Victor Pozsgay

Flavien Hirsch

Cyril Branciard

Nicolas Brunner

Covariance Bell Inequalities

We introduce Bell inequalities based on covariance, one of the most common measures of correlation. Explicit examples are discussed, and violations in quantum theory are demonstrated. A crucial feature of these covariance Bell inequalities is their nonlinearity; this has nontrivial consequences for the derivation of their local bound, which is not reached by deterministic local correlations. For our simplest inequality, we derive analytically tight bounds for both local and quantum correlations. An interesting application of covariance Bell inequalities is that they can act as "shared randomness witnesses": specifically, the value of the Bell expression gives device-independent lower bounds on both the dimension and the entropy of the shared random variable in a local model.

Bell inequalities limit the strength of possible correlations for any model satisfying a natural definition of locality formulated by Bell [1]. Thus, the violation of a Bell inequality indicates that no local model (in the sense of Bell) can reproduce the observed data, which is therefore said to be (Bell) nonlocal. This can be the case in quantum theory where, by performing well-chosen local measurements on an entangled quantum system, one may obtain nonlocal quantum correlations.

Beyond their fundamental interest, Bell inequalities are widely used tools, in particular in quantum information science [2]. Indeed, a Bell inequality violation represents a simple and strong test for certifying the presence of entanglement. Importantly this test is device-independent, in the sense that, assuming quantum theory, a Bell inequality violation implies the presence of entanglement without any assumption on the measuring devices, nor on the Hilbert space dimension. Moreover, the violation of a Bell inequality can also be used to certify the presence of genuine quantum randomness [3,4], or the security of a cryptographic key [5].

Consider two separate observers, performing local measurement on a shared physical system. The experiment results in some data, namely the joint conditional probabilities of observing a pair of outputs (measurement results) given a pair of inputs (choice of measurement settings). As they capture the strength of correlations in the data, it is intuitive that Bell inequalities are constructed based on some measure of correlation. For instance, the simplest and most famous Bell inequality of Clauser-Horne-Shimony-Holt (CHSH) [6] is a linear combination of simple correlation functions (see below). As these correlation functions are themselves linear combinations of the joint conditional probabilities, the CHSH inequality is a so-called linear Bell inequality. More generally, it is known that the set of distributions that are Bell local can be fully characterized via linear Bell inequalities [START_REF] Pitowsky | Quantum Probability-Quantum Logic[END_REF], and great efforts have been dedicated to find such inequalities, see e.g. Refs. [8][9][10][11][12][13].

It is however natural to ask if other measures of correlations can be used for devising Bell inequalities. Beyond the purely conceptual interest, this could be relevant in a practical context in which the full data is not available, but only certain specific (not necessarily linear) functions of the joint probabilities are. Moreover, it would be interesting to devise new types of Bell inequalities for the case of continuous measurement outcomes, for which the standard approach for constructing linear inequalities does not work anymore.

Several works have explored these ideas. First, entropic Bell inequalities [14][15][16] were obtained by considering the mutual information between the measurement outcomes. Later, inequalities based on higher moments of the distribution were also derived [17][18][19][20]. Finally, nonlinear Bell inequalities have been developed for discussing the generalization of Bell nonlocality to networks [21][22][23][24][25].

Here we introduce a novel class of nonlinear Bell inequalities based on the covariance of measurement results, a natural measure of correlations that is widely used in many different areas of science. We propose a general way to construct such Bell inequalities for covariances, and present, in the simplest case, a method for computing their local bound-a nontrivial problem due to the nonlinearity of the Bell expression. We show on a few explicit examples that our inequalities can detect quantum nonlocality. Moreover, we show that covariance Bell inequalities can be used to characterize shared randomness in a Bell test, providing device-independent lower bounds on the dimension and the entropy of the shared random variable in any local model.

Standard Bell inequalities.-Consider an experiment in which two parties, Alice and Bob, can both locally test different properties of a shared physical system. Let us label by x the measurement choice of Alice and y that of Bob, with the corresponding outcomes defining the random variables A x and B y , which take the values a and b, respectively. We will only consider non-signalling scenarios, in which the statistics of A x do not depend on the choice of y, and similarly the statistics of B y do not depend on x.

The statistics of the whole experiment, and thereby the correlations between Alice and Bob's measurement outcomes, are characterized by the joint conditional probability distribution P (a, b|x, y) = P (A x = a, B y = b). Such a distribution is called local (in the sense of Bell) if it admits a decomposition of the form P (a, b|x, y) = ρ(λ) P A (a|x, λ) P B (b|y, λ) dλ, (1) where λ represents the possible values a shared classical variable Λ (shared randomness), distributed with the density function ρ(λ). The local response functions, defined by the distributions P A (a|x, λ) and P B (b|y, λ) represent the local behavior of Alice and Bob's subsystems.

The set of local distributions-the "local set"-is constrained by linear Bell inequalities of the form B =

x,y,a,b

α ab|xy P (a, b|x, y) ≤ β, ( 2
)
with some real coefficients α ab|xy , and where the "local bound" β is the maximal value of the quantity B over all local distributions. For any given (finite) number of possible measurement settings and outcomes, the local set forms a polytope, and is thus fully characterized by a finite set of such linear Bell inequalities [START_REF] Pitowsky | Quantum Probability-Quantum Logic[END_REF].

As a concrete example, consider the case of two measurements for each party, labelled by x, y ∈ {0, 1}, with binary outcomes a, b ∈ {+1, -1}. Here we have the wellknown CHSH inequality [6] CHSH

≡ A 0 B 0 + A 0 B 1 + A 1 B 0 -A 1 B 1 ≤ 2, (3)
where the correlation functions A x B y are simply defined as the expectation values of the products of outcomes, A x B y = a,b a b P (a, b|x, y). Like many other linear Bell inequalities, the CHSH inequality can detect quantum nonlocal correlations, obtained by performing well chosen local measurements on a shared entangled state.

A Bell inequality for covariances.-Instead of using the correlation functions A x B y , other quantities can also be considered to characterize local distributions; indeed, Bell inequalities were for instance constructed for entropic quantities [14][15][16]. In this work we focus on another natural and widely used measure of correlation, namely the covariance. As mentioned previously, this may be of practical interest for situations where the full data is not available, and only covariances can be estimated. An advantage of using covariances is also that, like entropies, they are naturally defined for any number of possible measurement outcomes, and even for continuous outcomes. Unlike entropies however, they depend on the specific values given to the measurement outcomes; this allows one, in particular, to distinguish correlations versus anticorrelations.

The covariance of the two output variables A x and B y is defined as

cov(A x , B y ) = A x B y -A x B y , (4)
with A x B y defined as above, and similarly with A x = a,b a P (a, b|x, y) and B y = a,b b P (a, b|x, y) (for continuous values, the sums can simply be replaced by integrals, and the probabilities by probability density functions; note also that because of the non-signalling assumption, A x does not depend on y, and B y does not depend on x). We emphasize that because of the product term in Eq. ( 4), the covariance is a nonlinear function of the joint probabilities P (a, b|x, y).

We start by again considering the case of two measurements per party (x, y ∈ {0, 1}), but with now outcomes A x and B y that can take any values a, b in the interval [-1, +1]; this implies in particular that -1 ≤ cov(A x , B y ) ≤ 1 (note that if a, b are taken in any other bounded interval, they can, together with the corresponding covariance, simply be rescaled so that they lie in [-1, +1]). Our goal is to bound the set of possible values of the four covariances cov(A x , B y ) for local distributionsthe "local set" for covariances. To get some intuition, and inspired by the form of CHSH in Eq. ( 3), one may look for instance at the projection of this local set onto the 2-dimensional (covCHSH, covCHSH ) plane, with

covCHSH ≡ cov(A 0 , B 0 ) + cov(A 0 , B 1 ) + cov(A 1 , B 0 ) -cov(A 1 , B 1 ), (5)
and with covCHSH defined similarly, except that cov(A 0 , B 1 ) and cov(A 1 , B 0 ) come with minus signs. By optimizing the values of covCHSH ( ) numerically (but with reliable enough results), we obtained the local set represented in Fig. 1.

One immediately sees that contrary to the standard local polytope in the probability space, the local set in the covariance space is nonconvex, and that (unsurprisingly) it cannot be fully described by a finite number of Bell inequalities based on linear combinations of the covariances. Its full characterization thus looks much more complicated than that of the standard local polytope. Nevertheless, one can still derive some Bell inequalities for covariances that provide an outer approximation of the local set. As illustrated on Fig. 1, an example of such covariance Bell inequalities is the following:

covCHSH ≤ 16 7 . ( 6
)
In order to prove that this inequality indeed holds for any local distribution, it is in fact sufficient to restrict to distributions with binary outputs ±1. Those can be written as convex mixtures of finitely many deterministic (local) distributions, and one can then optimize the expression covCHSH over the corresponding weights. To take into account the constraint that these weights must be nonnegative and sum up to 1, one can introduce Karush-Kuhn-Tucker (KKT) multipliers [START_REF] Karush | Minima of Functions of Several Variables with Inequalities as Side Constraints[END_REF], which provide necessary conditions for a solution to be optimal. By considering decompositions onto different sets of deterministic distributions with nonzero weights, these KKT conditions simplify to a number of linear systems, which can easily be solved. The full details of the proof are given (together with an alternative approach) in Appendix A.

�������≤ �� � � � � � � ��� � �� 0 1 2 3
It is instructive to look more closely in the proof at decompositions with a given number d of deterministic distributions with nonzero weights. First note that for d = 1, i.e., for deterministic distributions, all covariances are 0, and therefore one also obtains covCHSH = 0. To obtain a nonzero value, one thus needs to consider mixtures of at least 2 distributions. For d = 2, one finds that the maximal possible value is covCHSH = 2, obtained for instance by the distribution P 2 = 1 2 (+ + / + -) + (--/ -+) (which also gives covCHSH = 2, see Fig. 1)-where the notation (A 0 A 1 /B 0 B 1 ) denotes a strategy in which Alice and Bob deterministically output A x and B y for inputs x and y, respectively. In order to reach the local upper bound in Eq. ( 6), one needs to go to d = 3; the distribution P Opt = 3 7 (+ + / + +) + 2 7 (-+ / --) + 2 7 (--/ -+), for instance, gives covCHSH = 16 7 (and covCHSH = 16 49 , see Fig. 1). One thus sees that reaching a given value of covCHSH requires mixing a certain number of deterministic strategies; this is an interesting feature, which will allow one to use the covariance inequality as a shared randomness witness (see below).

One may also look at possible violations of the covariance Bell inequality (6). One finds that it can indeed detect quantum nonlocality: that is, one can obtain covCHSH > 16 7 by performing local measurements on an entangled state. For example, the quantum correlations that reach the maximal quantum value (the "Tsirelson bound" [START_REF] Cirel'son | [END_REF]) of CHSH=2 √ 2 have vanishing marginals A x = B y = 0, and therefore also give covCHSH=2 √ 2 > 16 7 (point P Q on Fig. 1). In fact, this is the largest possible quantum violation, as one can prove that all quantum correlations must satisfy the inequality

covCHSH ≤ 2 √ 2 . ( 7
)
This is shown in Appendix B, where we also give a tighter characterization of the quantum set in terms of covariances. Note that in contrast to the local bound, marginals do not play any role in reaching the Tsirelson bound for covCHSH. Nevertheless, the nonlinearity of the Bell expression covCHSH also has interesting consequences in the quantum case. For instance, one can find two pure entangled states such that none of the states can individually violate the inequality (6), whereas a mixture of the two states can violate it; see Appendix C for details. Note, finally, that given the range of output variables, we have |cov(A x , B y )| ≤ 1, and therefore the algebraic maximum possible value of covCHSH is 4 (just like for CHSH). This value can be reached by the non-signalling distribution known as the Popescu-Rohrlich (PR) nonlocal box [28] (point P P R on Fig. 1), which again has vanishing marginals and also reaches the algebraic maximum of the CHSH expression.

Constructing other covariance Bell inequalities.-One can follow similar ideas to the ones developed above and derive other Bell inequalities based on covariances.

In the case of ternary inputs x, y ∈ {0, 1, 2} and binary outputs a, b ∈ {+1, -1} for instance, the local set in the space of expectation values A x B y is again fully characterized by CHSH-like inequalities; however, when considering the full probability space, i.e., including the marginals A x , B y , one gets a new inequivalent family of Bell inequalities of the form [9,10,29]

I 3322 ≡ A 0 B 0 + A 0 B 1 + A 0 B 2 + A 1 B 0 + A 1 B 1 -A 1 B 2 + A 2 B 0 -A 2 B 1 + A 0 + A 1 -B 0 -B 1 ≤ 4. ( 8
)
Inspired by the previous example of CHSH, one may look at the local bound when the expectation values A x B y are replaced by the covariances cov(A x , B y ). Marginal terms, viewed e.g. as A x = A x 1 1 B (where 1 1 B is the identity measurement operator for Bob, that always outputs b = 1) then simply drop out, as cov(A x , 1 1 B ) = 0.

We thus get the covariance Bell inequality (that allows for any outputs

a, b ∈ [-1, +1]) cov3322 ≡ cov(A 0 , B 0 ) + cov(A 0 , B 1 ) + cov(A 0 , B 2 ) + cov(A 1 , B 0 ) + cov(A 1 , B 1 ) -cov(A 1 , B 2 ) + cov(A 2 , B 0 ) -cov(A 2 , B 1 ) ≤ 9 2 . ( 9
)
Here the local bound was obtained through numerical optimization (up to machine precision); nevertheless, the reproducibility and reliability of the numerical results make us quite confident that the bound is correct. In Appendix D we present local distributions that reach it, and discuss quantum and super-quantum (non-signalling) violations.

More generally, one may consider constructing covariance Bell inequalities following the above recipe, starting from an arbitrary linear Bell inequality with binary outcomes. Note that instead of simply dropping the marginal terms, one could as well keep them, or replace them with other functions, e.g. with variances.

Generalising even further our approach for covariance Bell inequalities, one can also investigate Bell inequalities based on Pearson correlators, which can be thought of as normalised covariances. The Pearson correlator for two variables A x and B y with variances

σ(A x ) 2 = A 2 x -A x 2 and σ(B y ) 2 = B 2 y -B y 2 is defined as r(A x , B y ) = cov(Ax,By) σ(Ax)σ(By) (if either σ(A x ) or σ(B y ) is zero, we define r(A x , B y ) = 0)
. Following a similar construction as above, we get, for the case of binary inputs and now also restricting to binary outputs, the Bell inequality

rCHSH ≡ r(A 0 , B 0 ) + r(A 0 , B 1 ) + r(A 1 , B 0 ) -r(A 1 , B 1 ) ≤ 5 2 , ( 10
)
where the local bound was again obtained through numerical optimization. It must be emphasized that unlike for covariances, the local set for Pearson correlators is not the same when considering binary outputs or more possible outputs; for instance, with ternary outputs in {+1, 0, -1}, one can locally reach the value rCHSH = 2 √ 2, which turns out to also be the Tsirelson bound for rCHSH (whether we restrict to binary outputs or not), as proven already in Ref. [30] (we also give a proof of this in Appendix B). In Appendix D we present local distributions that reach these two local bounds of 5 2 and 2 √ 2, and discuss quantum and super-quantum violations of them. Of course, similar Bell inequalities with Pearson correlators could also be constructed, following a similar recipe as suggested above, starting e.g. from I 3322 or any other linear Bell inequality with binary outcomes.

Shared randomness witnesses.-Let us come back now to covCHSH. As emphasized before, the value of covCHSH that can be obtained locally depends on the number of deterministic strategies involved in the local strategy being used. This readily allows one to obtain a lower bound on the classical dimension d of the shared variable Λ (i.e., the number of different values it can take), or equivalently its max-entropy H max (Λ) ≡ log 2 d: from the discussion above, it follows that as soon as covCHSH > 0, one requires

��� � ��� (Λ) ��� �(Λ)
d ≥ 2, i.e. H max (Λ) ≥ 1; if covCHSH > 2, then d ≥ 3, i.e. H max (Λ) ≥ log 2 3.
One may also quantify the amount of shared randomness in terms of the Shannon entropy of Λ, H(Λ) ≡ λ q λ log 2 q λ (defined here, for simplicity, for the case of a discrete variable Λ that takes the value λ with probability P (Λ = λ) = q λ ). For a given value of covCHSH between 0 and 16 7 , one can also obtain a lower bound on H(Λ) by minimizing it for all local strategies. As it turns out, it suffices to optimise H(Λ) over decompositions onto deterministic strategies with binary outputs ±1-intuitively, local randomness does not help increase covCHSH. The details of our optimization are given in Appendix E, and the results we obtained are plotted on Fig. 2. Unsurprisingly, we find that min H(Λ) increases with the value of covCHSH. As soon as covCHSH > 0, one requires H(Λ) > 0; for covCHSH = 2, one must have H(Λ) ≥ 1 (recall that covCHSH = 2 can be reached by mixing two deterministic strategies with equal weights); finally, reaching the local bound of covCHSH = 16 7 requires at least as much shared randomness as in the distribution P Opt given previously, i.e.

H(Λ) ≥ -3 7 log 2 3 7 -2× 2 7 log 2 2 7
1.56. Thus covCHSH acts as a device-independent witness for characterizing the dimension and the entropy of the shared randomness of a local strategy in a Bell experiment. This complements recent works on device-independent tests of the dimension of quantum systems [31], and of the dimension [32] and entropy [33] of classical communications.

Discussion.-In this Letter we introduced Bell inequalities based on covariances, a natural measure of correlations, rather than on linear combinations of probabilities. We presented the simplest example of such a covariance Bell inequality with binary inputs, that echoes the well-known CHSH inequality, and investigated some of its properties. Proving analytically the local bound is not straightforward here, due to the nonlinearity of the Bell expression, but we could nevertheless provide a method to do so. Its quantum bound was also proven analytically.

Beyond this simplest example, we proposed a general recipe to construct, from any standard Bell inequality with binary outcomes, new Bell inequalities for covariances as well as for Pearson correlators. While our analytical method for proving the local bound generalizes in principle to any covariance Bell inequality obtained with our construction, in practice it becomes intractable for larger numbers of inputs. Nevertheless, for the examples we presented the local bounds could reliably be obtained numerically. It would be interesting to find a more efficient way to prove local bounds, as well as quantum bounds, for covariance-based Bell inequalities-or even to find other ways to characterize the local and quantum sets of admissible covariances (note that a tighter characterization of the quantum set than Eq. ( 7) can already be given, see Appendix B). One possible idea would be to look for some hierarchy of criteria-similar in spirit to that of Navascués-Pironio-Acín (NPA) for quantum correlations-that covariances must satisfy; indeed, one finds that covariances and Pearson correlators appear naturally in the NPA hierarchy [34].

An interesting feature of our covariance Bell inequalities is that they serve as a shared randomness witness. Indeed, in contrast with standard linear Bell inequalities where the local bound can always be attained by a single deterministic strategy, reaching the local bound-or in fact, any nontrivial value-for a Bell expression defined in terms of covariances requires, for a local strategy, to make use of shared randomness. We showed explicitly, on our simplest example, how the value of the Bell expression allows one to place device-independent lower bounds on the amount of shared randomness in a local model (both in terms of its dimension and its entropy). Such witnesses may help addressing certain problems in quantum nonlocality related to shared randomness [35][36][37], in particular finding what is the minimal amount of shared randomness necessary to simulate the correlations of entangled states admitting a local model [38]. Our example with binary inputs allows one to certify the use of relatively little shared randomness (e.g. a dimension of at least 3); it would be interesting to find (families of) covariance Bell inequalities that can certify larger amounts of shared randomness. While covariance Bell inequalities were found to provide a natural way to provide device-independent witnesses for shared randomness, such witnesses could also be studied in other frameworks, with other types of Bell inequalities that require shared randomness to be saturated-e.g. with entropic inequalities [14][15][16].

From a more practical perspective, our covariance Bell inequality could be useful in experimental situations where access to measurement data is limited, e.g. if only covariances can be measured, and where standard Bell inequalities cannot be used. This could be relevant to certain mesoscopic electronics experimental setups. With such applications in mind, it would also be interesting to also investigate similar covariance inequalities to the ones constructed here for other types of quantum nonlocality, e.g. for entanglement [39] or quantum steering [40] (we note that inequalities were constructed in these contexts for covariance matrices, although following very different approaches [41][42][43][44][45]).

Finally, Bell inequalities based on covariances could be useful for the study of nonlocal correlations in networks. When a network features several sources that are assumed to be independent, the set of local distributions (or rather, "N -local distributions" for a network with N independent sources) is typically nonconvex, and nonlinear Bell inequalities are necessary to give an effective description of the (N -)local set [21][22][23][24][25]. In fact, some recent work [46] has already made use of covariances in the context of quantum networks, however not constructing explicit Bell inequalities. The ideas developed here could also find some nice applications in providing ways to obtain new types of Bell inequalities, and possibly shared randomness witnesses, for networks. Obtaining the local bound for a standard (linear) Bell inequality is rather straightforward, as one can simply focus on deterministic strategies. However, such an approach does not work for covariance Bell inequalities because of their nonlinearity.

In this Appendix we give the details of the proof of the inequality (6), as sketched in the main text. We then present an alternative proof that also uses Karush-Kuhn-Tucker (KKT) multipliers and conditions [START_REF] Karush | Minima of Functions of Several Variables with Inequalities as Side Constraints[END_REF], although in a different way. While the former can in principle be adapted to any other covariance Bell inequality (but may not actually be practical because of the very high number of cases to consider for increased numbers of parameters), the latter exploits specific properties of the CHSH (2input / 2-output) scenario and uses fewer variables in the optimization; it is however less instructive with respect to the shared randomness required for obtaining a given value of covCHSH with a local model, and adapting it to another scenario will first require finding the right properties to exploit.

Optimizing over the weights of deterministic distributions

The proof here will be divided into three parts. The first one consists in showing that, without loss of generality, we can focus on a situation where only binary outputs are considered. The second part describes how standard methods for solving (quadratic) constrained problems can be applied. Finally, the third part shows how our quadratic problem can be simplified to a number of linear ones, and solved. x , B ( ) y ), are the same for the two local distributions P and P . 2Thus, any values of the covariances cov(A x , B y ) (for the various inputs x, y) obtained by a local distribution with inputs in [-1, +1] can also be obtained by a local distribution with inputs in {+1, -1} (and vice-versa, obviously). The local set for covariances is the same in both cases, and to characterize it it thus suffices to restrict to local distributions with binary outputs ±1.

b. KKT conditions

Our goal is now to find the largest possible value of the quantity covCHSH for any local strategy with binary inputs and outputs. Such a local distribution can be represented by a 16-dimensional vector P , the components of which are the 16 joint probabilities P (a, b|x, y) 3 -and which in fact, because of the normalization and nonsignaling constraints, live in a space of dimension only dim P = 8. 4 Moreover, P admits a convex decomposition in terms of the vertices of the local polytope of the form P = k q k P det k with q k ≥ 0 , k q k = 1, and where the vectors P det k represent the 16 deterministic local strategies. (Note already that due to Carathéodory's theorem [47], for any local distribution P it will in fact be sufficient to consider at most dim P + 1 = 9 deterministic strategies in the decomposition.) Evaluating covCHSH, we get

covCHSH({q k } k ) = k q k C kk - i,j q i q j C ij , (A1) where C ij = A i 0 B j 0 + A i 0 B j 1 + A i 1 B j 0 -A i 1 B j 1 = ±2 and A i
x refers to Alice's output for the measurement x given that strategy P det i is used, and similarly for B j y -so that in particular, C kk = CHSH( P det k ) is the CHSH value obtained for the deterministic strategy P det k . We shall now maximize covCHSH({q k } k ) with respect to the variables q k -which, because of the second term in Eq. (A1), is a nonlinear optimization problem. To account for the constraints that q k ≥ 0 and k q k = 1, we introduce KKT multipliers (i.e., Lagrange-like multipliers) λ k and µ, and define the following Lagrangian:

L = covCHSH({q k } k ) + k λ k q k + µ (1- k q k ) = k q k C k - i,j q i q j Z ij + k λ k q k + µ (1- k q k ). (A2)
Given the regularity of the problem, with the constraints being affine functions of the variables, the (local) maxima must then necessarily fulfill the KKT conditions [START_REF] Karush | Minima of Functions of Several Variables with Inequalities as Side Constraints[END_REF]:

∂L ∂q k = C kk - i q i (C ik + C ki ) + λ k -µ = 0 ∀k, ∂L ∂µ = 1 - k q k = 0, λ k q k = 0, λ k ≥ 0, q k ≥ 0. (A3) c.

Reduction to linear systems of equations and resolution

Because of the constraint λ k q k = 0, the above system of equations is nonlinear. One can however reduce it to a number of different linear systems of equations by imposing that a given subset of weigths q k take nonzero values, and by considering separately the case of each possible subset. For those q k in the given subset under consideration, the last line of Eq. (A3) simply becomes λ k = 0, q k > 0, while for the other q k 's, it becomes q k = 0, λ k ≥ 0-with in both cases, λ k = -C kk + i q i (C ik + C ki ) + µ according to the first line of Eq. (A3). For a given number d of nonzero weights, the system (A3) then reduces to d + 1 linear equality constraints and 16 (also linear) inequality constraints (with d strict and 16 -d nonstrict inequalities) for the remaining d + 1 nontrivial variables q k and µ.

Note that if only d = 1 weight q k is nonzero (and therefore equal to 1), then the distribution P (a, b|x, y) is deterministic, which implies that all covariances-and hence covCHSH-are zero. Furthermore, as already mentioned previously, due to Carathéodory's theorem it is sufficient to only consider subsets with at most d = 9 (out of 16) nonzero q k 's-indeed, if a maximum is reached by some decomposition involving more than 9 deterministic distributions, then it is also reached by another decomposition giving the same distribution, but involving no more than 9 deterministic distributions. We used Mathematica to solve, for each value of d between 2 and 9, the 16 d different linear systems of equations corresponding to all the different subsets containing d nonzero weights (which makes a total of 9 d=2 16 d = 50 626 linear systems to consider). More specifically, we first found the solutions that satisfy all equality constraints and then checked which ones also satisfy the inequality constraints. To obtain the local bound of our covariance Bell inequality (6), there then remains to check what is the largest possible value of covCHSH for all these feasible solutions.

and covCHSH = ± 16 49 ). The distributions in the table are grouped by pairs (brackets on the left), in which the two decompositions are obtained from one another by flipping all outputs.

For d = 2, all 16 2 = 120 systems of equations have solutions that satisfy all equality constraints (out of which, 56 have unique solutions and 64 are still underdetermined), but only 4 out of them have (unique) solutions that also satisfy the inequality constraints. Those 4 solutions all give covCHSH = 2, which is the maximal value that one can reach by mixing only 2 local deterministic distributions.

With d = 3 nonzero weights q k , again all 16 3 = 560 systems of equations have solutions that satisfy all equality constraints (432 with unique solutions and 128 still underdetermined), but only 8 of them have (unique) solutions that also satisfy the inequality constraints. Those solutions are listed in Table I; they all have the similar form P = 3 7 P det k1 + 2 7 P det k2 + 2 7 P det k3 , and all give the value covCHSH = 16 7 . As it turns out, this value of 16 7 obtained by suitably mixing 3 deterministic local distributions cannot be increased by mixing more distributions. For d = 4 for instance, we find that out of the 16 4 = 1 820 systems of equations, only 1 516 have solutions that satisfy all equality constraints, out of which only 14 have solutions that also satisfy the inequality constraints. 6 of these (still underdetermined) solutions give a local maximum of covCHSH = 2, while the other 8 (unique) solutionsall of the form P = found for any of the corresponding systems of KKT conditions. As one can see, the maximal value of covCHSH obtained for all cases is thus 16 7 ; this concludes the proof that this is indeed the value of the local bound in our covariance Bell inequality (6).

It should be clarified that although Table II gives local maxima of 2 or 9 4 for d ≥ 4, this does not mean that those are the maximum values of covCHSH when mixing d ≥ 4 deterministic local distributions. Indeed, the table gives local maxima obtained in a set delimited by strict inequalities q k > 0, while the suprema over these sets may be obtained for some q k → 0. In fact, covCHSH can get arbitrarily close to 16 7 with a mixture of d ≥ 4 deterministic local distributions by mixing an optimal decomposition with 3 distributions from Table I, with a tiny amount of d -3 other distributions.

Another observation of interest is that all solutions of the KKT conditions that give a local maximum value of covCHSH = 2 (for = 2, 4, 6, 8, as listed in the table) are of the general form P = q 1 (+ + / + +) + (--/ --) + q 2 (+ + / + -) + (--/ -+) + q 3 (+ -/ + +) + (-+ / --) + q 4 (+ -/ -+) + (-+ / + -) with q 1 , q 2 , q 3 , q 4 ≥ 0 and 2(q 1 + q 2 + q 3 + q 4 ) = 1: for d = 2, three of the coefficients q k are zero (which, for each remaining nonzero q k , gives a unique solution to the corresponding system of equations); for d = 4, two of the coefficients q k are zero; for d = 6, one of the coefficients q k is zero; and for d = 8, all coefficients q k are nonzero (which indeed makes the solutions in these last 3 cases underdetermined). As one can see, in this general decomposition each deterministic distribution comes with its "opposite", in which all outputs are flipped, with the same weight. This implies in particular that all marginal expectation values A x and B y are zero, that the covariances cov(A x , B y ) are equal to the expectation values A x B y , and therefore that the expression of covCHSH simply reduces to that of CHSH (i.e., the nonlinear part vanishes)-which is indeed bounded by 2 for local distributions. This highlights the crucial role played by the nonlinear terms A x B y in covCHSH: these are precisely the terms that allow the local bound of covCHSH to be greater than that of CHSH; and in order to reach a value of covCHSH larger than 2 with a local distribution, one needs at least one pair of settings x, y for which A x B y = 0.

Optimizing over the expectation values AxBy , Ax and By

We now present a second possible approach for the proof of Eq. ( 6), which builds on the observation that the local bound 16 7 on covCHSH follows from just imposing the local bound 2 on CHSH (which must be respected by any local distribution), and that probabilities are nonnegative. More precisely, as previously we note that one can restrict to binary outputs a, b = ±1; as it turns out, it will then suffice to impose, together with CHSH ≤ 2, that P (a, b|x, y) ≥ 0 for all x, y, a, b satisfying ab = (-1) xy+1 . Instead of writing covCHSH as a function of the weights of deterministic distributions in a local decomposition, here we will view it directly as a function of the expectation values A x B y , A x and B y . Our goal is thus to maximize the value of covCHSH under the above constraints-which can also all be expressed in terms of those expectation values, by writing in particular P (a, b|x, y) = 1 4 (1 + a A x + b B y + ab A x B y ). As previously, let us introduce KKT multipliers λ and λ xyab (for the 8 combinations of x, y, a, b such that ab = (-1) xy+1 ), and define the Lagrangian

L({ A x B y , A x , B y } xy , λ, {λ xyab } xyab ) = covCHSH + λ (2 -CHSH) + x,y,a,b: ab=(-1) xy+1 λ xyab P (a, b|x, y). (A4)
The KKT necessary conditions for optimality are Table III. Optimal solutions of the KKT conditions of Eq. (A5) (together with λ = 5 7 in all cases and λ xyab as given in Eq. (A6)), giving covCHSH = 16 7 and defining the same distributions as in Table I.

∂L ∂ A x B y = (-1) xy (1 -λ) + 1 4 a,b λ xyab ab = 0 ∀x, y, ∂L ∂ A x = - y (-1) xy B y + 1 4 y,a,b λ xyab a = 0 ∀x, ∂L ∂ B y = - x (-1) xy A x + 1 4 x,a,b λ xyab b = 0 ∀y, λ (2 -CHSH) = 0, λ ≥ 0, 2 -CHSH ≥ 0, λ xyab P (a, b|x, y) = 0, λ xyab ≥ 0, P (a, b|x, y) ≥ 0, ∀x, y, a, b : ab = (-1) xy+1 . (A5) A0B0 A0B1 A1B0 A1B1 A0 A1 B0 B1
The first 3 lines define 8 linear equality constraints for the 8 multipliers λ xyab , which can be solved and give

λ xyab = 2 1 -λ -a(-1) y A x -b(-1) x B ȳ , (A6)
with x = 1 -x and ȳ = 1 -y. We are then left with just the last couple of lines of Eq. (A5), so that either λ = 0 or CHSH = 2, and for each x, y, a, b such that ab = (-1) xy+1 , either λ xyab = 0 or P (a, b|x, y) = 0. We can then consider the 2 1+8 = 512 corresponding cases separately, and solve for each case the resulting linear system of 9 equations for the 9 variables A x B y , A x , B y and λ.

Using Mathematica we found feasible solutions for those equations in 320 of the 512 cases. However, one still needs to check if these solutions satisfy the remaining inequality constraints in Eq. (A5); only in 207 of those cases could we find solutions that indeed satisfy all KKT conditions. Among all these solutions, we found that the maximal value of covCHSH was 16 7 , which again proves that this is the local bound of our covariance Bell inequality (6). This value of 16 7 was obtained for the 8 solutions listed in Table III, which indeed define the same 8 distributions as those listed (in the same order) in Table I.

Generalizing our proofs to other covariance Bell inequalities

The two versions of the proof of inequality ( 6) presented above could in principle be generalized to calculate the local bound of any covariance Bell inequality constructed from a linear Bell inequality with binary outcomes, following the recipe we suggest in the main text. However, as the number of inputs, and therefore the dimension of the relevant probability space, increase, the number of different cases to consider in the proof may become too large to be tractable in a reasonable time.

For the case of cov3322 with ternary inputs for instance, after reducing to binary outputs and taking into account normalization and non-signalling constraints, the relevant probability space is of dimension dim P = 15. Following the first approach for the proof above, local distributions can be decomposed onto the 2 6 deterministic local distributions, and the number of different cases to consider (corresponding to the number of subsets of d nonzero weights q k , with 2 ≤ d ≤ dim P + 1) is

16 d=2 2 6
d ≥ 7 × 10 14 , a number far too large for all cases to be considered separately.

One may hope that following our second approach for the proof, instead, may reduce the number of cases to be considered down to a tractable one. The first thing to check here would be which constraints (locality constraints in terms of CHSH or I 3322 inequalities, together with nonnegativity of the probabilities constraints) are sufficient to impose in order to obtain the local bound on cov3322. We did not follow this approach any further, and leave as an open question, whether this would be tractable enough to provide an analytical proof of inequality (9).

Instead, to obtain the local bound of inequality ( 9), we resorted to numerical optimization, by optimizing over the 2 6 = 64 weights q k in a local decomposition of P = k q k P det k . Although the optimization problem is nonconvex and involves quite a few free parameters, we found that the result of the numerical optimization was stable enough when starting from different starting points, which makes us confident that the local bound of 9 2 in inequality ( 9) is indeed correct.

We note, finally, that the same proof techniques as above do not work for Bell inequalities with Pearson correlators like inequality (10). The first point to be noticed is that, as emphasized before, the local sets for binary outputs a, b ∈ {+1, -1} and for more outputs a, b ∈ [-1, +1] are not the same-indeed we found different local bounds for binary and ternary outputs. Even restricting to a fixed number of inputs and trying to follow our approach with KKT multipliers, the Lagrangians we would write in either of the two proof versions would not be nicely quadratic functions of the weights q k in a local decomposition of the form P = q k P det k (as in Eq. (A2)), or of the expectation values A x B y , A x and B y (as in Eq. (A4)), so that the KKT necessary conditions for linearity would not simplify to linear systems of equations. Instead of pursuing such an approach, we again resorted here to numerical optimizations. As before, the numerical results we obtained were stable enough to make us confident that the local bounds we give are correct.

inspired by that of Landau (for rCHSH) in Ref. [30].

α x |β y = ψ| ( Âx ⊗1 B -Âx ψ 1 AB )(1 A ⊗ By -By ψ 1 AB ) |ψ = ψ| Âx ⊗ By |ψ -Âx ψ By ψ = A x B y -A x B y = cov(A x , B y ) (B1)
and similarly,

α x |α x = A 2 x -A x 2 = σ(A x ) 2 , β y |β y = B 2 y -B y 2 = σ(B y ) 2 . (B2)
Restricting to measurement operators with eigenvalues in [-1, +1],5 this implies in particular that

α x | = α x |α x = σ(A x ) ≤ 1,
β y | = β y |β y = σ(B y ) ≤ 1. (B3)
We thus obtain covCHSH

= α 0 |β 0 + α 0 |β 1 + α 1 |β 0 -α 1 |β 1 ≤ α 0 | |β 0 + |β 1 + α 1 | |β 0 -|β 1 ≤ |β 0 + |β 1 + |β 0 -|β 1 ≤ 2 |β 0 + |β 1 2 + |β 0 -|β 1 2 = 2 |β 0 2 + |β 1 2 ≤ 2 √ 2, ( B4
)
which is necessarily satisfied by any quantum correlation. This bound can be reached by adequate measurements on a maximally entangled state, see Appendix B 3 below.

Let us mention here that following further the proof of Ref. [30], one can actually also refine the characterization of the quantum set in terms of covariances. Indeed, defining the ket vectors |γ 0 = |α 0 , |γ 1 = |α 1 , |γ 2 = |β 0 , |γ 3 = |β 1 and the matrix Γ with coefficients Γ ij = γ i |γ j (for 0 ≤ i, j, ≤ 3), one necessarily has that Γ is positive semidefinite (as Γ = M † M with M = i |γ i i|, where {|i } is the computational basis). That is, taking Eqs. (B1)-(B2) into account, one must have

Γ =      σ(A 0 ) 2 Γ 01 cov(A 0 ,B 0 ) cov(A 0 ,B 1 ) Γ * 01 σ(A 1 ) 2 cov(A 1 ,B 0 ) cov(A 1 ,B 1 ) cov(A 0 ,B 0 ) cov(A 1 ,B 0 ) σ(B 0 ) 2 Γ 23 cov(A 0 ,B 1 ) cov(A 1 ,B 1 ) Γ * 23 σ(B 1 ) 2      ≥ 0. (B5)
Given that the variances are upper-bounded by 1, one also has Γ ≥ 0, with the matrix Γ obtained from Γ above by replacing all diagonal terms by 1. Now, it can be shown (following the same arguments as in Ref. [30], see also Ref. [34]) that the condition that there exists (real or complex) coefficients Γ 01 and Γ 23 such that Γ ≥ 0 requires arcsin cov(A 0 , B 0 ) + arcsin cov(A 0 , B 1 )

+ arcsin cov(A 1 , B 0 ) -arcsin cov(A 1 , B 1 ) ≤ π (B6)
(as well as the symmetrical inequalities obtained by permuting the measurement inputs and/or flipping the sign of the measurement outcomes), which must therefore necessarily be satisfied by quantum correlations, and strengthens the condition that covCHSH ≤ 2 √ 2.

Tsirelson bound for rCHSH

To obtain the Tsirelson bound for rCHSH, we follow a very similar approach to the one above, considering now normalized ket vectors |α x = |α x /|||α x || and βy = |β y /|||β y || (if |α x or |β y are null vectors, we define |α x or βy to be any unit vector orthogonal to all other ones). With these, we now have αx βy = r(A x , B y ).

(B7)

Because of the normalization of |α x , βy and r(A x , B y ) (via the division by the variances σ(A x ), σ(B y )), we do not need to restrict here to measurement operators with eigenvalues in [-1, +1] (or such that σ(A x ), σ(B y ) ≤ 1). With the same calculations as in Eq. (B4), we obtain, for any choice of (bounded) measurement operators Âx , By , rCHSH = α0 β0 + α0 β1 + α1 β0 -α1 β1

≤ • • • ≤ 2 √ 2. ( B8
)
This bound can again be reached quantum mechanically, see below.

As just mentioned, here no restriction is imposed on the measurement outcomes (provided only that σ(A x ), σ(B y ) < ∞). The bound above also holds for local distributions (which can always be realized quantum mechanically) when measurement outcomes are not necessarily restricted to be binary; as we show explicitly in Appendix D 2 b, it can also be reached locally with ternary outcomes. It is quite remarkable that the local and the Tsirelson bounds coincide in this case.

Let us finally mention again that, as it was done in Ref. [30] (see also Ref. [34]), a tighter characterization of the quantum set in terms of Pearson correlators can also be given: any quantum correlation satisfies arcsin r(A 0 , B 0 ) + arcsin r(A 0 , B 1 ) + arcsin r(A 1 , B 0 ) -arcsin r(A 1 , B 1 ) ≤ π (B9) (as well as the symmetrical inequalities). (6) and (10) The Tsirelson bounds for covCHSH and rCHSH can be reached by the same quantum mechanical correlations as those that reach the maximal quantum value of CHSH = 2 √ 2. For instance, consider that Alice and Bob share a two-qubit maximally entangled state |φ + = 1 Appendix C: Higher quantum violations of Ineq. (6) with mixed states: a consequence of nonlinearity

Explicit quantum violations of inequalities

In this Appendix we discuss an example illustrating the fact that the nonlinearity of the Bell inequality (6) has nontrivial consequences also in the quantum case. Specifically, consider two pure entangled states of the form

|φ θ = cos θ 2 |0, 0 + sin θ 2 |1, 1 , (C1) |ψ θ = sin θ 2 |0, 0 + cos θ 2 |1, 1 , (C2)
with θ ∈ ]0, π/2], and let us then define the equal mixture of these states: We numerically estimated the largest value of the Bell expression ( 5), optimizing over all local qubit measurements (including positive-operator valued measures) by Alice and Bob, for various values of the parameter θ. The result is presented in Fig. 3. We found, up to numerical precision, that for the states |φ θ and |ψ θ the value of covCHSH was maximal when using the Pauli measurements Â0 = σx , Â1 = σy , B0 = σx-σy √ 2 and B1 = σx+σy √ 2 , giving covCHSH = 2 √ 2 sin θ (orange dashed curve on Fig. 3), while the optimal measurements on the state ρ θ were found to be Â0 = σz , Â1 = σx , B0 = σz+sin θ σx √ 1+sin 2 θ and B1 = σz-sin θ σx √ 1+sin 2 θ , giving covCHSH = 2 1 + sin 2 θ (blue curve on Fig. 3). This gives violations of inequality (6) for θ > arcsin 16/7 2 √ 2 0.94 in the first case, and for θ > arcsin ( 16/7 2 ) 2 -1 0.59 in the second case. We thus find that in the range of parameters 0.59 θ 0.94, none of the two pure states can individually violate the covariance Bell inequality (6), whereas the mixture (C3) does violate it. This effect is possible only via the nonlinearity of the covariance Bell inequality.

ρ θ = 1 2 (|φ θ φ θ | + |φ θ φ θ |) . (C3) ρ θ ϕ θ � ψ θ  π 2 0
Note that the pure states |φ θ and |ψ θ and the mixture ρ give in fact the same correlation functions A x B y for any (projective) Pauli measurements Âx = σAx and By = σBy for Alice and Bob. The difference resides in the marginals A x and B y , which are opposite for |φ θ and |ψ θ and are therefore unbiased for the equal mixture ρ θ .

Let us finally mention that we also investigated quantum violations of inequality (6) using mixed entangled states of rank 3 and higher (up to dimension 6 × 6). We always found that the optimal violations were obtained with pure or rank-2 states. It thus seems that mixing more than 2 pure entangled states does not help increase the violations. σz + sin θ 2 σx , Â1 = B1 = 1 -sin 2 θ 4 σzsin θ 2 σx , Â2 = B2 = σx ). These violate inequality (9) for θ > arcsin 9/2 5 1.12 and θ > arcsin 9 2 -4 = π 4 , respectively. As it was the case with covCHSH, we thus find a range of parameters (for π 4 < θ 1.12) for which neither |φ θ nor |ψ θ violates inequality (9), but their equal mixture does-which is a consequence of the nonlinearity of cov3322. Similarly to covCHSH this property does not seem to extend to mixtures of more than two pure states, that is, the violation of rank-3 states (or higher) does not appear to be larger than when mixing only two pure states. It is also interesting to note that within the families under consideration, cov3322 is violated by a smaller range of states than covCHSH. This is in contrast to the relationship between the standard CHSH and I 3322 linear Bell inequalities, where all entangled (possibly mixed) quantum states violating (a symmetry of) the former also violate the latter [10]. 6 Note, finally, that the algebraic maximum value of cov3322 = 8 can be attained by the non-signalling distributions such that

A 0 B 0 = A 0 B 1 = A 0 B 2 = A 1 B 0 = A 1 B 1 = -A 0 B 2 = A 2 B 0 = -A 2 B 1 = 1 (with A 2 B 2
remaining a free parameter) and with all marginals A x = B y = 0. Such distributions also reach the maximal non-signalling value of I 3322 = 8. For the case of binary outputs (taking e.g. a, b = ±1) we find, again numerically, that the local bound for rCHSH is 5 2 , as in inequality (10). This bound can be reached by the 8 distributions that decompose as mixtures of the same triplets of deterministic distributions P det k1 , P det k2 , P det k3 as those in Table I of Appendix A (i.e., those that allow one to reach the local bound of the covariance inequality (6)), but taking here uniform mixtures, P = As emphasized before, in the case of Pearson correlators, considering binary outputs or a larger alphabet of possible outputs does make a difference. As shown in Appendix B 2, the value of rCHSH, for any choice of possible outputs, that can be reached locally is upper-bounded by 2 √ 2. As it turns out, this bound can be reached with ternary outputs a, b = +1, 0, -1, for instance by the local distribution P = 4 9 (+ + / + 0) + 4 9 (+ -/0+) + 1 9 (00/ --). Note that if any output variable A x or B y is deterministic, its variance is zero and by convention we took r(A x , B y ) = 0. This implies that rCHSH defined in Eq. ( 10) contains at most 2 nonzero Pearson correlators, which implies that rCHSH ≤ 2. Furthermore, if the two output variables of any party are both deterministic, then rCHSH = 0. Nevertheless, one can also locally reach a value of rCHSH arbitrarily close to the bound 2 √ 2 by certain distributions that get arbitrarily close to some deterministic ones: e.g., P = (1 -)(+ + / + 0) + 2 (+ -/ -+) + 2 (-+ / --) with 0 < < 1 gives rCHSH = 2(1 + √ 1 -)/ √ 2 -→ 2 √ 2 as → 0 (note that Alice still has a binary output ±1 here; only Bob has a ternary output +1, 0, -1, for one of his inputs only). Hence, in this case the value of rCHSH cannot be used as a shared randomness witness.

c. Violations of inequality (10) As proven in Appendix B, the Tsirelson bound for rCHSH is 2 √ 2, which can be reached by adequate (binary) projective measurements on a maximally 2-qubit entangled states. The quantum correlations thus obtained thus violate the inequality (10) (with the local bound 5 2 , for the case where one restricts to binary outputs), but do not violate the analogous inequality for arbitrary outcomes, when the local bound also becomes 2 √ 2 (see above).

We note that for rCHSH, the nonlinearity of the Bell expression does not seem to exhibit the same interesting features as covCHSH and cov3322 presented in the previous appendices: the maximal value of a mixture of pure states appears, from our numerical investigations, to always be attained by one of the pure states.

Finally, note that algebraic maximum of rCHSH=4 can be obtained by a PR box (just as for CHSH and covCHSH).

Figure 2 . 16 7 2 .

 2162 Figure 2. Minimum Shannon entropy H(Λ) (blue curve) and max-entropy Hmax(Λ) = log 2 d (orange dashed curve) of the shared variable Λ that is required to obtain any given value of the Bell expression covCHSH with a local model. For instance, obtaining covCHSH = 2 requires at least one bit of (Shannon or max-) entropy, while reaching the maximum local value covCHSH = 16 7 2.29 requires H(Λ) 1.56 and Hmax(Λ) ≥ log 2 3 1.58, i.e. d ≥ 3.

-a 2 ( 1 -b 2 )a 2 1+b b 2 P

 2122 a. Reduction to binary outputs Consider a local distribution P (a, b|x, y) with outputs a, b ∈ [-1, +1]. 1 Let then Alice and Bob post-process their outputs locally, in the following way: they replace each output a (b) by a = +1 (b = +1) with probability 1+a 2 ( 1+b 2 ), or by a = -1 (b = -1) with probability 1, thus defining a new local distribution P (a , b |x, y) = ab 1+a (a, b|x, y) with binary outputs a , b ∈ {+1, -1}. One can easily check that the expectation values A

1 .

 1 Tsirelson bound for covCHSHFirst note that by Neumark's dilation theorem, any quantum correlation can be obtained by projective measurements on pure states in Hilbert spaces of large enough dimensions.Consider a bipartite state |ψ ∈ H AB = H A ⊗ H B and projective measurement operators Âx for Alice and By for Bob (with the random variables A x , B y corresponding to the results of those measurements). Let us then define the ket vectors|α x = ( Âx ⊗ 1 B -Âx ψ 1 AB ) |ψ and |β y = (1 A ⊗ By -By ψ 1 AB ) |ψ ,where 1 X denotes the identity operator acting on the Hilbert space H X and where Âx ψ = ψ| Âx ⊗ 1 B |ψ and By ψ = ψ| 1 A ⊗ By |ψ . With these definitions, we have

√ 2 ( 2 .

 22 |0, 0 + |1, 1 ) and perform the local measurements Â0 = σz , Â1 = σx , B0 = σz+σx √ 2 and B1 = σz-σx √ 2 , where σz and σx denote the Pauli matrices. In that case the marginal expectation values A x and B y vanish, so that covCHSH and rCHSH simply reduce to CHSH, and we indeed obtain covCHSH = rCHSH = CHSH = 2 √ By then rotating for instance Bob's measurement settings together around the y axis of the Bloch sphere, one obtains the full circle drawn on Fig. 1.

Figure 3 .

 3 Figure 3. Maximal value of the Bell expression covCHSH for the pure entangled states |φ θ and |ψ θ (dashed orange curve), and for the mixture ρ θ (solid blue curve). The horizontal dotted red line represents the local bound of the Bell inequality (6). In the interval θ ∈ ] 0.59, 0.94], none of the two pure states violates the inequality, while the mixture does.

2 .

 2 Bell inequality (10) for Pearson correlators a. Optimal local strategies with binary outputs a, b = ±1

1 36

 1 k3 . E.g., for the first one: P = (+ + / + +) + 1 3 (-+ / --) + 1 3 (--/ -+). This can indeed be seen as follows: taking Â2 = B0 = -1 1, the expression of cov3322 in Eq. (8) simplifies toI 3322 = A 0 B 1 + A 0 B 2 + A 1 B 1 -A 1 B 2 + 2 =CHSH 01/12 + 2, where CHSH 01/12 is a symmetry of CHSH, Eq. (3). Clearly if there exist measurement operators giving CHSH 01/12 > 2, one then also gets a value I 3322 > 4. Note that the same argument does not extend to covCHSH and cov3322: for the choice of Â2 = B0 = -1 1, we just get cov3322 = covCHSH 01/12 ≤ 4, so that no violation of inequality (9) is possible with those measurement settings. b. Optimal local strategies with ternary outputs a, b = +1, 0, -1

Table II .

 II For d = 6 and 8, we also get local maxima of covCHSH = 2, while for d = 5, 7 and 9, no solutions are

	d	16 d	# consistent systems of equalities	# consistent systems incl. inequalities	local max covCHSH
	2	120	120	4	2
	3	560	560	8	16 7
	4	1 820	1 516	14	2 / 9 4
	5	4 368	3 376	0	-
	6	8 008	1 896	4	2
	7	11 440	688	0	-
	8	12 870	154	1	2
	9	11 440	16	0	-
					3 8 P det k1 + 3 8 P det k2 + 1 8 P det k3 + 1 8 P det k4
					(one solution being obtained for instance by taking
					P det k1 = (++/++), P det k2 = (-+/--), P det k3 = (+-/-+)
					and P det k4 = (--/ -+))-give covCHSH = 9 4 . The res-
					ults for all values of d from 2 to 9 are summarized in

Table II .

 II For any number d between 2 and 9 (1st column), the 2nd column of the table gives the number of subsets of deterministic distributions P k with nonzero weigths q k , which defines the number of linear systems-obtained from the KKT conditions (A3)-to consider in the proof. The numbers of systems which have solutions when only considering the equality constraints, or when also including the inequality constraints, are given in the next two columns. The last column gives the local maxima of covCHSH obtained from these feasible solutions of the KKT conditions.

Note that given the translational invariance of covariances (i.e., cov(Ax, By) = cov(Ax + α, By + β) for any real values α, β), it would be equivalent to take a, b in any interval of length

Similarly in the case of binary outputs, instead of ±1 one could take any two real values with a difference of 2.

We note that the restriction to bounded operators with eigenvalues in [-1, +1] is in fact not necessary in the proof, as it is sufficient to just impose σ(Ax), σ(By) ≤ 1 (for any state |ψ ) for the Tsirelson bound on covCHSH to follow.

Acknowledgements.-We thank Alastair A. Abbott, Joe Bowles, Michael J. W. Hall and Paul Skrzypczyk for discussions. We acknowledge financial support from the Swiss National Science Foundation (Starting grant DIAQ and QSIT) and from the French National Research Agency ('Retour Post-Doctorants' program ANR-13-PDOC-0026).

Table I. Optimal decompositions of the form P = 3 7 P det k 1 + 2 7 P det k 2 + 2 7 P det k 3 obtained as solutions of the KKT conditions of Eq. (A3), giving covCHSH = 16 7 . The first 4 distributions give covCHSH = 16 49 and thus project onto the point POpt on Fig. 1 (while the last 4 give covCHSH = - 16 49 ; note also that there exist 8 other similar decompositions that give covCHSH = -16

Appendix B: Tsirelson bounds and quantum violations for covCHSH and rCHSH

In this Appendix we give the proofs that the Tsirelson bounds for both covCHSH and rCHSH are the same as for CHSH, namely 2 √ 2. The proofs presented here are Appendix D: Optimal local strategies and violations for our covariance Bell inequalities (9)- (10) In this Appendix we present some local strategies that reach the local bounds of our Bell inequalities ( 9) and (10), and show how these can be violated quantum mechanically and by non-signalling correlations. Recall that the local bounds were obtained numerically; however, we are confident that the numerics are trustworthy enough, and therefore that the local strategies we present below are indeed optimal.

Covariance Bell inequality (9) a. Optimal local strategies

To obtain the local bound for cov3322 we ran a numerical optimization several times and conclude from the stability of the result that it must be 9 2 , as in inequality (9).

There are 4 different local distributions that give the value cov3322 = 9 2 , and that decompose onto three deterministic correlations with weights 3 8 , 3 8 and 1 4 : namely, P = 3 8 (+++/+++)+ 3 8 (--+/--+)+ 1 4 (-+-/-+-), P = 3 8 (+++/++-)+ 3 8 (--+/---)+ 1 4 (+--/-++) (using here the notation (A 0 A 1 A 2 /B 0 B 1 B 2 )), and the other 2 "opposite decompositions", where all outputs are flipped.

b. Violations of inequality (9) Inequality ( 9) can be violated by quantum correlations. For example, taking a two-qubit maximally entangled state |φ + = 1 √ 2 (|0, 0 + |1, 1 ) and suitable measurements (see below), one can obtain cov3322= 5, which is also the highest value for I 3322 (Eq. ( 8)) obtainable with qubits and projective measurements [9,10,29] (note that in that case, all marginal terms vanish and cov3322 effectively reduces to I 3322 ). We have performed an extensive numerical search, considering entangled states of dimension up to 5 × 5, and could not find any larger quantum violation than cov3322=5. It thus remains an open question, whether higher dimensional entangled states could lead to larger violations, as it is the case for the standard I 3322 inequality [48].

Considering again the families of states |φ θ , |ψ θ and ρ θ of Eqs. (C1)-(C3), the largest values of cov3322 we found (numerically, up to machine precision) are cov3322 = 5 sin θ for |φ θ or |ψ θ (obtained e.g. for Â0 = B1 = σx , Â1 = B2 = cos π 3 σx + sin π 3 σy , Â2 = -B0 = cos 2π 3 σx + sin 2π 3 σy ) and cov3322 = 4 + sin 2 θ for ρ θ (obtained e.g. for Â0 = B0 = 1 -sin 2 θ

Appendix E: Minimal amount of shared randomness required to reach a given value of covCHSH

In this appendix we provide the details on how one can find the minimal amount of shared randomness-quantified here in terms of the Shannon entropy H(Λ)required to locally reach a given value of covCHSH.

A general local model provides a decomposition for the probabilities P (a, b|x, y) as in Eq. (1)

which clearly satisfies minH det Λ (c) ≥ minH gen Λ (c). We will then prove that these two functions in fact coincide: to reach the minimal amount of shared randomness for a given value of covCHSH, it is indeed enough to just consider binary outputs ±1, and decompositions onto deterministic local response functions.

Local decompositions onto deterministic response functions with binary outputs ±1

Let us first note that for binary inputs and binary outputs, there is only a finite number (16) of local deterministic strategies for Alice and Bob together. Even though the local decomposition (1) with deterministic response functions may involve more than 16 different values for λ (and even continuous values, in an appropriate limit), clearly the entropy H(Λ) is minimized when the different values of λ that define the same deterministic strategies are grouped together. This implies that it is sufficient to consider here local decompositions of the form P = k q k P det k onto the 16 different deterministic distributions P det k . The optimization problem in Eq. (E2) then reduces to optimize the 16 weights q k under the constraints q k ≥ 0, k q k = 1 and covCHSH({q

We performed this optimization numerically, for various values of c = covCHSH. Our results are shown on Fig. 2 (the function minH det Λ coincides with min H(Λ) shown there). We repeated the numerical optimization several times, starting with different random starting points; the consistency of the results we obtained make us confident that we indeed reached the global minima in Eq. (E2).

What we found (up to numerical precision) is that for c = covCHSH ≤ 2, the minimum in (E2) is reached by mixing only 2 deterministic local distributions with nonzero weights: for instance, taking P = q(+ + / + +) + (1-q)(--/ --) gives covCHSH = 8q(1-q), H(Λ) = H({q, 1-q}) = -q log 2 q -(1-q) log 2 (1-q), and leads

2.29, we found that the minimum in (E2) is reached by mixing now 3 deterministic local distributions with nonzero weights: for instance, with distributions of the form P = q 1 (+ + / + +) + q 2 (-+ / --) + q 3 (--/ -+). That gives covCHSH = (1 -q 1 )(1 + 7q 1 ) -(q 2 -q 3 ) 2 , H(Λ) = H({q 1 , q 2 , q 3 }). For a given value of covCHSH = c, that reduces to optimizing over just one parameter, say q 1 , as q 2 and q 3 are then given by q 2,3 =

(which requires

, so that all q k are between 0 and 1). For c 2.27 we find that the optimal is obtained for q 1 = 3+ √ 16-7c 7

(and q 2 = q 3 = 1-q1 2 ); for c 2.27 we cannot give an analytical solution (however, a good approximation seems to be q 1 1 2 -0.01(c -2)). We note that as can clearly be seen on Fig. 2, minH det Λ is a (strictly) increasing function of c = covCHSH.

General local decompositions

For a general local decomposition of the form (1), one gets

with A x λ = a a P (a|x, λ) and B y λ = b b P (b|y, λ) (so that A x = ρ(λ) A x λ dλ and similarly for B y ).

Looking at this expression, it is clear that for a given distribution ρ(λ) (and therefore a given value of H(Λ)) and some given response functions of Bob, the maximal value of covCHSH is obtained when | A 0 λ | = 1, i.e., when Alice uses deterministic response functions with output ±1 (specifically, Alice's optimal response function is

A similar reasoning can be followed for Bob. We then conclude that for a given value of H(Λ), the maximal value of covCHSH is obtained when Alice and Bob use deterministic response functions with outputs a, b = ±1: essentially, local randomness (or considering other possible outputs in [-1, +1]) does not help increase the value of covCHSH.

Note, however, that the optimization problem in Eq. (E1) is subtly different: rather than maximizing covCHSH for a given value of H(Λ), we want to mimimize H(Λ) for a given value of covCHSH. To conclude here we will use the previous observation, that minH det Λ (c) is an increasing function of c = covCHSH.

Consider indeed a (general) local decomposition of a distribution P with a given amount of shared randomness H(Λ) = h, and which gives some value covCHSH = c. From the above reasoning, it follows that there exists another distribution P using the same shared random variable Λ (hence, with the same value H(Λ) = h) but now decomposed onto deterministic local response functions with binary outputs ±1, that gives a value covCHSH = c ≥ c. This then implies that minH det Λ (c ) ≤ h and, because minH det Λ is an increasing function, minH det Λ (c) ≤ h as well. Hence, any general local decomposition that gives some value covCHSH = c necessarily satisfies H(Λ) ≥ minH det Λ (c). This implies that minH gen Λ (c) ≥ minH det Λ (c)-and therefore, minH gen Λ (c) = minH det Λ (c)with minH det Λ (c) estimated above.


S Badoux
email: sven.badoux@usherbrooke.ca

S A A Afshar

B Michon

A Ouellet

S Fortier

D Leboeuf

T P Croft

C Lester

S M Hayden

H Takagi

K Yamada

D Graf

N Doiron-Leyraud

Louis Taillefer
email: louis.taillefer@usherbrooke.ca

Critical doping for the onset of Fermi-surface reconstruction by charge-density-wave order in the cuprate superconductor La 2-x Sr x CuO 4

The Seebeck coefficient S of the cuprate superconductor La2-xSrxCuO4 (LSCO) was measured in magnetic fields large enough to access the normal state at low temperatures, for a range of Sr concentrations from x = 0.07 to x = 0.15. For x = 0.11, 0.12, 0.125 and 0.13, S/T decreases upon cooling to become negative at low temperatures. The same behavior is observed in the Hall coefficient RH(T ). In analogy with other hole-doped cuprates at similar hole concentrations p, the negative S and RH show that the Fermi surface of LSCO undergoes a reconstruction caused by the onset of charge-density-wave modulations. Such modulations have indeed been detected in LSCO by X-ray diffraction in precisely the same doping range. Our data show that in LSCO this Fermisurface reconstruction is confined to 0.085 < p < 0.15. We argue that in the field-induced normal state of LSCO, charge-density-wave order ends at a critical doping pCDW = 0.15 ± 0.005, well below the pseudogap critical doping p ⋆ ≃ 0.19.

Since the discovery of quantum oscillations [START_REF] Doiron-Leyraud | Quantum oscillations and the Fermi surface in an underdoped high-Tc superconductor[END_REF] and a negative Hall coefficient R H [START_REF] Leboeuf | Electron pockets in the Fermi surface of hole-doped high-Tc superconductors[END_REF] in the cuprate superconductor YBa 2 Cu 3 O y (YBCO), it has become clear that the Fermi surface of underdoped YBCO undergoes a reconstruction at low temperature that produces a small electron pocket [START_REF] Taillefer | Fermi surface reconstruction in high-Tc superconductors[END_REF], in a doping range from p = 0.08 to p ≃ 0.15 [START_REF] Leboeuf | Lifshitz critical point in the cuprate superconductor YBa2Cu3Oy from high-field Hall effect measurements[END_REF]. This Fermi-surface reconstruction (FSR) was also detected as a sign change in the Seebeck coefficient S(T ), going from positive at high temperature to negative at low temperature [START_REF] Chang | Nernst and Seebeck Coefficients of the Cuprate Superconductor YBa2Cu3O6.67: A Study of Fermi Surface Reconstruction[END_REF]. A strikingly similar change of sign in S(T ) observed in the cuprate La 1.8-x Eu 0.2 Sr x CuO 4 (Eu-LSCO) [START_REF] Laliberté | Fermi-surface reconstruction by stripe order in cuprate superconductors[END_REF] suggested that the stripe order known to exist in Eu-LSCO [START_REF] Fink | Phase diagram of charge order in La1.8-xEu0.2SrxCuO4 from resonant soft x-ray diffraction[END_REF] -a combination of charge-density-wave (CDW) and spin-densitywave (SDW) modulations -is responsible for the FSR in both materials. The observation of CDW modulations in YBCO by NMR [START_REF] Wu | Magnetic-field-induced charge-stripe order in the high-temperature superconductor YBa2Cu3Oy[END_REF] and X-ray diffraction (XRD) [START_REF] Ghiringhelli | Long-Range Incommensurate Charge Fluctuations in (Y,Nd)Ba2Cu3O6+x[END_REF][START_REF] Chang | Direct observation of competition between superconductivity and charge density wave order in YBa2Cu3O6.67[END_REF] confirmed this conjecture, and demonstrated that it is the CDW (and not the SDW) modulations that cause the FSR.

In YBCO, the drop in R H (T ) and S/T begins at a temperature T max that peaks at p = 0.12 (Fig. 1a). The drop is attributed to the CDW modulations detected by XRD [START_REF] Hücker | Competing charge, spin, and superconducting orders in underdoped YBa2Cu3Oy[END_REF][START_REF] Blanco-Canosa | Resonant x-ray scattering study of charge-density wave correlations in YBa2Cu3O6+x[END_REF] and NMR [START_REF] Wu | Incipient charge order observed by NMR in the normal state of YBa2Cu3Oy[END_REF] below a temperature T CDW in the same doping range as the FSR [START_REF] Leboeuf | Lifshitz critical point in the cuprate superconductor YBa2Cu3Oy from high-field Hall effect measurements[END_REF], with T CDW also peaking at p = 0.12 (Fig. 1a).

In HgBa 2 CuO 4+δ (Hg1201), high-field measurements of Hall and Seebeck coefficients revealed a similar FSR [START_REF] Doiron-Leyraud | Hall, Seebeck, and Nernst Coefficients of Underdoped HgBa2CuO 4+δ : Fermi-Surface Reconstruction in an Archetypal Cuprate Superconductor[END_REF], confirmed by the observation of quantum oscillations [START_REF] Barišić | Universal quantum oscillations in the underdoped cuprate superconductors[END_REF] and again attributed to XRD-detected CDW modulations [START_REF] Tabis | Charge order and its connection with Fermi-liquid charge transport in a pristine high-Tc cuprate[END_REF]. All this suggests that CDW modulations and the associated FSR are generic properties of hole-doped cuprates in the vicinity of p = 0.12. A major outstanding question is : Up to what critical doping p CDW do CDW modulations extend in the phase diagram (Fig. 1), in particular in the field-induced normal state at T = 0 ? In this context, the material LSCO offers a powerful platform, since good crystals can be grown with p up to 0.3 and beyond. CDW modulations have been observed in LSCO with XRD, at p ≃ 0.12 [START_REF] Croft | Charge density wave fluctuations in La2-xSrxCuO4 and their competition with superconductivity[END_REF][START_REF] Thampy | Rotated stripe order and its competition with superconductivity in La1.88Sr0.12CuO4[END_REF][START_REF] Christensen | Bulk charge stripe order competing with superconductivity in La2-xSrxCuO4 (x=0.12)[END_REF], but there is little information about the associated FSR.

In this Article, we report high-field measurements of the Seebeck coefficient in LSCO single crystals at several dopings, which show that S becomes negative in the normal state at low temperature in precisely the doping range where CDW modulations are detected by XRD. R H is also found to be negative in that range. The FSR in LSCO is therefore very similar to the FSR in YBCO and Hg1201. Our data show that the FSR does not extend above p = 0.15, strong evidence that CDW order in LSCO ends at a critical doping p CDW = 0.15. This implies that in the normal state of LSCO the phase of CDW order ends well before the pseudogap phase, which ends at the critical doping p ⋆ ≃ 0. 19 (ref. 20).

Methods.-Single crystals of LSCO were grown by the flux-zone technique with Sr concentrations x = 0.085, 0.11, 0.12 and 0.13 at the University of Bristol, x = 0.07 and 0.125 at the University of Tokyo, x = 0.144 and 0.15 Charge density-wave (CDW) modulations are detected by Xray diffraction below TCDW (green triangles) in YBCO (up triangles [START_REF] Hücker | Competing charge, spin, and superconducting orders in underdoped YBa2Cu3Oy[END_REF], down triangles [START_REF] Blanco-Canosa | Resonant x-ray scattering study of charge-density wave correlations in YBa2Cu3O6+x[END_REF]) and LSCO (up triangles [START_REF] Croft | Charge density wave fluctuations in La2-xSrxCuO4 and their competition with superconductivity[END_REF], down triangle [START_REF] Christensen | Bulk charge stripe order competing with superconductivity in La2-xSrxCuO4 (x=0.12)[END_REF]). Spin-density-wave (SDW) modulations are detected by neutron diffraction below TSDW (blue squares) in YBCO [START_REF] Haug | Neutron scattering study of the magnetic phase diagram of underdoped YBa2Cu3O6+x[END_REF] and LSCO [START_REF] Chang | Tuning competing orders in La2-xSrxCuO4 cuprate superconductors by the application of an external magnetic field[END_REF][START_REF] Kofu | Hidden Quantum Spin-Gap State in the Static Stripe Phase of High-Temperature La2-xSrxCuO4 Superconductors[END_REF][START_REF] Wakimoto | Observation of incommensurate magnetic correlations at the lower critical concentration for superconductivity in La2-xSrxCuO4 (x=0.05)[END_REF][START_REF] Lake | Antiferromagnetic order induced by an applied magnetic field in a hightemperature superconductor[END_REF][START_REF] Kimura | Neutron-scattering study of static antiferromagnetic correlations in La2-xSrxCu1-yZnyO4[END_REF]. When plotted as S/T vs T , the normal-state Seebeck coefficient peaks at a temperature Tmax (full red circles) before it drops at low temperature due to Fermi-surface reconstruction (YBCO, ref. [START_REF] Laliberté | Fermi-surface reconstruction by stripe order in cuprate superconductors[END_REF]; LSCO, this work, Figs. 3 and4). A similar Tmax can also be defined for the Hall coefficient (open red circles), below which RH(T ) drops at low temperature (YBCO,ref. 4).

at Tohoku University. Samples were cut in the shape of rectangular platelets, with typical dimensions 0.5 mm × 1.0 mm × 0.1 mm. The hole concentration (doping) p is taken to be p = x. The (zero-resistance) superconducting transition temperature of the 8 samples is T c = 12.7, 20.2, 26.2, 27.5, 28.0, 32.3, 37.2, and 36.5 K for p = 0.07, 0.085, 0.11, 0.12, 0.125, 0.13, 0.144, and 0.15, respectively. The

(c) (e) (d) (f) Figure 2.
Isotherms of the Seebeck coefficient in LSCO, plotted as S/T vs magnetic field H, at various temperatures, as indicated, for six samples, with x = 0.07 (a), x = 0.085 (b), x = 0.125 (c), x = 0.13 (d), x = 0.144 (e), and x = 0.15 (f). For x = 0.125 and 0.13, S/T at high H decreases at low temperature, to reach negative values. For x = 0.144, S/T also decreases at low temperature, below 15 K. This decrease is the signature of FSR. In contrast, for x = 0.07 and 0.15, S/T at the highest measured field keeps increasing with decreasing temperature down to the lowest temperature. This shows that there is no FSR at those dopings, at least down to 4 K and 9 K, respectively. The same is true at x = 0.085, at least down to 15 K.

Seebeck coefficient was measured, as described elsewhere [START_REF] Laliberté | Fermi-surface reconstruction by stripe order in cuprate superconductors[END_REF], at Sherbrooke (all samples) up to H = 20 T, at the National High Magnetic Field Laboratory (NHMFL) in Tallahassee up to H = 34 T (x = 0.125 and 0.15) and up to H = 45 T (x = 0.13), and at the Laboratoire National des Champs Magnétiques Intenses (LNCMI) in Grenoble up to H = 34 T (x = 0.07 and 0.144). The Hall coefficient of samples with x = 0.11, 0.12, 0.125 and 0.13 was measured, as described elsewhere [START_REF] Leboeuf | Lifshitz critical point in the cuprate superconductor YBa2Cu3Oy from high-field Hall effect measurements[END_REF], at Sherbrooke in H = 16 T. All crystals have an orthorhombic crystal structure and they are twinned. The thermal gradient or electrical current was applied in the basal plane, while the magnetic field was applied along the c axis.

Seebeck coefficient.-In Fig. 2, the Seebeck data for 6 samples are plotted as S/T vs H for several tempera- All data points represent the normal state, for which the solid lines are a guide to the eye, except the lowest point for each of x = 0.144 and x = 0.15 (panel b). For these two points, the isotherms are still going up the superconducting transition (Fig. 2). The dashed lines are an extension of the normal-state behavior based on extrapolating those isotherms beyond 34 T. Tmax marks the temperature below which S/T decreases at low temperature (arrow), in some cases to reach negative values, as seen here for x = 0.125. This decrease is the signature of Fermi-surface reconstruction (FSR). Note how the data for x = 0.144 and x = 0.15 split below T ≃ 30 K, with the former dropping at low T due to FSR and the latter showing no decrease, and hence no FSR (at least down to 9 K). tures. We see that for x = 0.125 (Fig. 2c) and x = 0.13 (Fig. 2d), S becomes negative at high field and low temperature. This shows that a negative S is a property of the normal state of LSCO at these dopings, as in YBCO, Eu-LSCO and Hg1201. At x = 0.144, we see that at high field S/T decreases when the temperature drops below T = 15 K (Fig. 2e). In contrast, no such decrease is observed at x = 0.15, down to the lowest tempera- Comparison of LSCO (red) and YBCO (green) at p = 0.12. a) Temperature dependence of the X-ray intensity associated with the CDW modulations, normalized at Tc, detected in LSCO [START_REF] Croft | Charge density wave fluctuations in La2-xSrxCuO4 and their competition with superconductivity[END_REF] and YBCO [START_REF] Chang | Direct observation of competition between superconductivity and charge density wave order in YBa2Cu3O6.67[END_REF]. Lines are a guide to the eye.

The cusp is at Tc. b) Normal-state Seebeck coefficient of LSCO (this work) and YBCO [START_REF] Laliberté | Fermi-surface reconstruction by stripe order in cuprate superconductors[END_REF], measured in a magnetic field as indicated, plotted as S/T vs T . Tmax is the temperature below which S/T drops to reach negative values at low temperature (arrow), the signature of Fermi-surface reconstruction (FSR). This Tmax is plotted as full circles in Fig. 1. Lines are a guide to the eye. c) Hall coefficient of LSCO at H = 16 T and YBCO at H = 15 T [START_REF] Leboeuf | Electron pockets in the Fermi surface of hole-doped high-Tc superconductors[END_REF], plotted as eRH/V , where e is the electron charge and V the volume per planar Cu atom. Tmax is the temperature below which RH(T ) drops to reach negative values at low temperature (arrow), another signature of FSR. Tmax is plotted as open circles in Fig. 1a [4].

ture (Fig. 2f). At x = 0.07, S/T increases steadily with decreasing T at high field, down to the lowest temperature (Fig. 2a). This is also true at x = 0.085 (Fig. 2b). Although here our data only goes to 20 T, the crossing of the lowest isotherms shows that S/T keeps increasing down to T = 15 K, at least. In Figs. 3 and4, we plot S/T vs T , at high field. In Fig. 3a, we see that the drop in S/T at x = 0.125 to negative values starts below a temperature T max ≃ 40 K. This is also the case at x = 0.11 and 0.13 (Fig. 4a). In Fig. 4b, we compare data on 3 samples taken in identical conditions, at H = 16 T. (Although the LSCO sample with x = 0.12 was only measured up to 18 T, S/T at T = 8 K is increasingly negative with increasing H (inset of Fig. 4b), confirming that a negative S is a property of the normal state also at that doping.) The location of the peak in S/T vs T is seen to decrease from T max = 45 K at x = 0.12, to T max = 42.5 K at x = 0.125, to T max = 40 K at x = 0.13. Those T max values are plotted on the phase diagram of LSCO in Fig. 1b. Raising the doping further, we observe that T max continues its steady descent. Indeed, at p = 0.144, S/T now peaks at T max ≃ 15 K (Fig. 3b). Extrapolating this trend yields T max → 0 at p → 0.15 (Fig. 1b). Our data at x = 0.15 confirm this, with S/T showing no decrease down to at least 9 K (Figs. 2f and3b). This shows that FSR in LSCO ends at a critical doping p FSR = 0.15 ± 0.005.

At x = 0.07, the normal-state S/T increases monotonically with decreasing T , down to our lowest temperature (Fig. 3a). There is clearly no FSR at that doping. At x = 0.085, although we only measured up to 18 or 20 T, we observe that S/T at H = 18 T increases as T → 0, at least down to 15 K (Fig. 2b). So here T max < 15 K. In Fig. 1b, we plot T max vs p for our 8 samples, with their uncertainty, and thereby delineate the region where FSR occurs in the Tp phase diagram of LSCO. We see that the FSR region peaks at p ≃ 0.12 and is confined between p ≃ 0.085 and p = p FSR = 0.15 ± 0.005.

Hall coefficient.-In Fig. 5c, the Hall coefficient of our LSCO crystal with x = 0.12, measured at H = 16 T, is plotted as R H vs T . We see that R H (T ) drops below T ≃ 50 K and becomes negative below T ≃ 20 K. Data for our crystals with x = 0.11, 0.125 and 0.13 are very similar, also negative at low T , all in excellent agreement with prior low-field data on single crystals of LSCO with x = 0.12 [START_REF] Suzuki | Hall coefficient of La1.88-yYySr0.12CuO4 (y=0,0.04) at low temperatures under high magnetic fields[END_REF]. (The absence of a negative R H in previous high-field data on thin films of LSCO [START_REF] Balakirev | Quantum Phase Transition in the Magnetic-Field-Induced Normal State of Optimum-Doped High-Tc Cuprate Superconductors at Low Temperatures[END_REF] may be due to the higher disorder of such samples.) A similar drop in R H (T ) has been seen in Eu-LSCO [START_REF] Cyr-Choinière | Enhancement of the Nernst effect by stripe order in a high-Tc superconductor[END_REF] and in La 1.4-x Nd 0.6 Sr x CuO 4 (Nd-LSCO) [START_REF] Noda | Evidence for One-Dimensional Charge Transport in La2-x-yNdySrxCuO4[END_REF], when p ≃ 0.12; in both materials, it is closely linked to the onset of CDW order.

Discussion.-Taken together, the negative Hall and Seebeck coefficients in the normal state of LSCO are conclusive evidence of FSR in this material, in the vicinity of p = 0.12. This adds to the previous three cases, namely YBCO, Eu-LSCO and Hg1201. In all 4 cases, the FSR occurs in a region of the T -p phase diagram where CDW modulations have been detected by XRD (Fig. 1). The link between CDW and FSR is robust.

It is instructive to compare LSCO and YBCO. The two phase diagrams are similar (Fig. 1). In both cases, T max and T CDW peak at p = 0.12, and the region of FSR is confined to similar ranges -from p ≃ 0.085 to p = 0.15 in LSCO and from p = 0.08 to p ≃ 0.15 in YBCO [START_REF] Leboeuf | Lifshitz critical point in the cuprate superconductor YBa2Cu3Oy from high-field Hall effect measurements[END_REF]. In Fig. 5, we compare data for LSCO and YBCO directly, at p = 0.12. The CDW modulations detected by XRD emerge below a temperature twice as high in YBCO com-pared to LSCO (Fig. 5a) : T CDW ≃ 150 K in YBCO vs T CDW ≃ 75 K in LSCO. Correspondingly, the FSR is detected at a temperature twice as high in YBCO compared to LSCO, with T max ≃ 100 K in YBCO vs T max ≃ 50 K in LSCO (Fig. 5b). All this suggests that CDW ordering is a stronger tendency in YBCO than in LSCO. Intriguingly, the superconducting transition temperature T c is roughly twice as high in YBCO as compared to LSCO (see cusp in Fig. 5a). This raises the interesting possibility that the same underlying mechanism, perhaps magnetic, fuels both superconductivity and CDW order [START_REF] Efetov | Pseudogap state near a quantum critical point[END_REF].

Given that FSR in LSCO ends at p FSR = 0.15, we infer that this is also where CDW order ends. This is consistent with recent XRD measurements that detect no CDW modulations in LSCO at x = 0.15 [32]. (The same consistency is observed at x = 0.085, where again no CDW modulations are detected by XRD [32].) We thus arrive at a key information : the CDW phase in LSCO ends at the critical doping p CDW = 0.15. This is distinctly below the critical point where the pseudogap phase is believed to end in LSCO, at p ⋆ ≃ 0.19, as determined from the normal-state resistivity measured in high magnetic fields [START_REF] Cooper | Anomalous Criticality in the Electrical Resistivity of La2-xSrxCuO4[END_REF]. This clear separation reveals that the pseudogap phase is not caused by the CDW ordering. Instead, it suggests that CDW order is a secondary instability of the pseudogap phase. A very similar separation was recently observed in YBCO from high-field Hall effect measurements, with p CDW = 0.16 ± 0.005 and p ⋆ ≃ 0.19 [START_REF] Badoux | Change of carrier density at the pseudogap critical point of a cuprate superconductor[END_REF]. This strongly suggests that a separation of p CDW and p ⋆ is a generic property of cuprates.

Summary.-Our high-field measurements of the See-beck coefficient in the cuprate superconductor LSCO reveal that its normal-state Fermi surface undergoes a reconstruction at low temperature, in the doping range 0.085 < p < 0.15. In analogy with the cuprates YBCO, Eu-LSCO and Hg1201, we attribute this FSR to the CDW modulations detected by XRD in the very same doping range. Combined with XRD data on LSCO, our Seebeck data make a compelling case that CDW modulations disappear at p = p CDW = 0.15, so that the field-induced non-superconducting ground state of LSCO above p = 0.15 has no CDW order. Because the pseudogap phase in the normal state of LSCO extends up to p ≃ 0.19, we infer that the pseudogap is not tied to CDW ordering. Instead, the CDW modulations appear to be a secondary instability of the pseudogap phase.

Figure 1 .

 1 Figure 1. Temperature-doping phase diagram of the cuprate superconductors YBCO (a) and LSCO (b). The superconducting transition temperature Tc is drawn as a black line.Charge density-wave (CDW) modulations are detected by Xray diffraction below TCDW (green triangles) in YBCO (up triangles[START_REF] Hücker | Competing charge, spin, and superconducting orders in underdoped YBa2Cu3Oy[END_REF], down triangles[START_REF] Blanco-Canosa | Resonant x-ray scattering study of charge-density wave correlations in YBa2Cu3O6+x[END_REF]) and LSCO (up triangles[START_REF] Croft | Charge density wave fluctuations in La2-xSrxCuO4 and their competition with superconductivity[END_REF], down triangle[START_REF] Christensen | Bulk charge stripe order competing with superconductivity in La2-xSrxCuO4 (x=0.12)[END_REF]). Spin-density-wave (SDW) modulations are detected by neutron diffraction below TSDW (blue squares) in YBCO[START_REF] Haug | Neutron scattering study of the magnetic phase diagram of underdoped YBa2Cu3O6+x[END_REF] and LSCO[START_REF] Chang | Tuning competing orders in La2-xSrxCuO4 cuprate superconductors by the application of an external magnetic field[END_REF][START_REF] Kofu | Hidden Quantum Spin-Gap State in the Static Stripe Phase of High-Temperature La2-xSrxCuO4 Superconductors[END_REF][START_REF] Wakimoto | Observation of incommensurate magnetic correlations at the lower critical concentration for superconductivity in La2-xSrxCuO4 (x=0.05)[END_REF][START_REF] Lake | Antiferromagnetic order induced by an applied magnetic field in a hightemperature superconductor[END_REF][START_REF] Kimura | Neutron-scattering study of static antiferromagnetic correlations in La2-xSrxCu1-yZnyO4[END_REF]. When plotted as S/T vs T , the normal-state Seebeck coefficient peaks at a temperature Tmax (full red circles) before it drops at low temperature due to Fermi-surface reconstruction (YBCO, ref.[START_REF] Laliberté | Fermi-surface reconstruction by stripe order in cuprate superconductors[END_REF]; LSCO, this work, Figs.3 and 4). A similar Tmax can also be defined for the Hall coefficient (open red circles), below which RH(T ) drops at low temperature(YBCO, ref. 4).

Figure 3 .

 3 Figure 3. Seebeck coefficient of LSCO, plotted as S/T vs temperature T , measured in a magnetic field H = 0 (open circles), 16 T (full circles) and 34 T (squares), for four samples, with x = 0.07 and 0.125 (a), and x = 0.144 and 0.15 (b). The data in panel b are normalized to their value at T = 100 K.All data points represent the normal state, for which the solid lines are a guide to the eye, except the lowest point for each of x = 0.144 and x = 0.15 (panel b). For these two points, the isotherms are still going up the superconducting transition (Fig.2). The dashed lines are an extension of the normal-state behavior based on extrapolating those isotherms beyond 34 T. Tmax marks the temperature below which S/T decreases at low temperature (arrow), in some cases to reach negative values, as seen here for x = 0.125. This decrease is the signature of Fermi-surface reconstruction (FSR). Note how the data for x = 0.144 and x = 0.15 split below T ≃ 30 K, with the former dropping at low T due to FSR and the latter showing no decrease, and hence no FSR (at least down to 9 K).

Figure 4 .

 4 Figure 4. Same as in Fig.3, for samples with x = 0.11 (yellow), x = 0.12 (blue), x = 0.125 (red) and x = 0.13 (green), measured at H = 16 T (full circles), 17.5 T (open squares) and 44 T (full squares). The data in panel b are normalized to their value at T = 100 K. FSR is clearly observed in all four samples, as a drop in S/T at low temperature. Inset of panel b: Isotherm at T = 8 K for x = 0.12, showing that S/T becomes increasingly negative with increasing field, demonstrating that the negative S is a property of the normal state.

Figure 5 .

 5 Figure 5. Comparison of LSCO (red) and YBCO (green) at p = 0.12. a) Temperature dependence of the X-ray intensity associated with the CDW modulations, normalized at Tc, detected in LSCO[START_REF] Croft | Charge density wave fluctuations in La2-xSrxCuO4 and their competition with superconductivity[END_REF] and YBCO[START_REF] Chang | Direct observation of competition between superconductivity and charge density wave order in YBa2Cu3O6.67[END_REF]. Lines are a guide to the eye. The cusp is at Tc. b) Normal-state Seebeck coefficient of LSCO (this work) and YBCO[START_REF] Laliberté | Fermi-surface reconstruction by stripe order in cuprate superconductors[END_REF], measured in a magnetic field as indicated, plotted as S/T vs T . Tmax is the temperature below which S/T drops to reach negative values at low temperature (arrow), the signature of Fermi-surface reconstruction (FSR). This Tmax is plotted as full circles in Fig.1. Lines are a guide to the eye. c) Hall coefficient of LSCO at H = 16 T and YBCO at H = 15 T[START_REF] Leboeuf | Electron pockets in the Fermi surface of hole-doped high-Tc superconductors[END_REF], plotted as eRH/V , where e is the electron charge and V the volume per planar Cu atom. Tmax is the temperature below which RH(T ) drops to reach negative values at low temperature (arrow), another signature of FSR. Tmax is plotted as open circles in Fig.1a[START_REF] Leboeuf | Lifshitz critical point in the cuprate superconductor YBa2Cu3Oy from high-field Hall effect measurements[END_REF].

Nature 531, 210 (2016).

Acknowledgements.-We thank J. Chang, N. E. Hussey, M.-H. Julien, P. A. Lee, B. J. Ramshaw, S. Sachdev, J. L. Tallon, and G. A. Sawatzky for stimulating discussions. A portion of this work was performed at the National High Magnetic Field Laboratory, which is supported by the National Science Foundation Cooperative Agreement No. DMR-1157490, the State of Florida, and the U.S. Department of Energy. Another portion of this work was performed at the Laboratoire National des Champs Magnétiques Intenses of the CNRS, member of the European Magnetic Field Laboratory. L.T. acknowledges support from the Canadian Institute for Advanced Research (CIFAR) and funding from the National Science and Engineering Research Council of Canada (NSERC), the Fonds de recherche du Québec -Nature et Technologies (FRQNT), the Canada Foundation for Innovation (CFI) and a Canada Research Chair.


domain: sde phys

text:
Daniel Hallatt
email: daniel.hallatt@univ-lille.fr

Hugues Leroux

Flavie Emmanuel Dubois

Flavie Braud

Exploring the response of Ryugu-inspired phyllosilicates to a pulsed laser

  L'archive ouverte pluridisciplinaire

Introduction: Thanks to JAXA's Hayabusa2 mission, samples returned from Ryugu mark a new era of opportunity in extraterrestrial materials science [1]. Along with NASA's ongoing sample return mission to Bennu, returning samples from C-group asteroids promises to offer novel insights into the material history of our Solar System. One material that may draw particular attention in the returned samples is a family of sheet-structured minerals called phyllosilicates. Phyllosilicates are not only believed to occur in widespread abundance on both sampled asteroids (thanks to consistent and recognizable spectral signatures [2,3]), but, thanks to their alterable mineralogy, may be able to be used as an indicator of the environment that they (and potential parent bodies which they comprised) have experienced. It is from an understanding of their response to the environment of space, coupled with the relative geologically inactivity of primitive asteroidal hosts, such as Ryugu, that the use of phyllosilicates as tracers of extraterrestrial-material processing may make it possible to understand the role of asteroids as messengers of times past.

In the context of asteroids, materials are thought to experience a diverse series of phenomena depending on the dynamic history of the body. These may include thermal alteration (radiogenic decay, solar heating, etc.), irradiation (solar-wind, comic-rays, etc.), aqueous alteration, and impact (micrometeorite bombardment, large scale collisions, etc.). It is with regards to the latter phenomena that we present, as a partial derivation from a broader investigation (a doctoral thesis), work using a pulsed laser to simulate the effect of impact on phyllosilicates. In its full scope, the entirety of the study also includes a bulk look into the behaviour of phyllosilicates under thermal conditions, followed by 2stage light-gas gun experiments to compare to these laser-simulated impacts.

Methods: The Ryugu-relevant phyllosilicates reported here include a pressed pellet of powdered serpentine, and polished chips of cut serpentine and saponite. To validate our experimental conditions, polished chips of San Carlos olivine were subject to identical laser conditions as the phyllosilicate targets, to be exploited as a reference to the more mature literature on lasered olivine. Each sample was irradiated with a femtosecond pulsed laser source (Tangerine from Amplitude Systèmes), embedded within an Oxford Lasers machining. The 1030 nm laser had a unit pulse duration and energy of ~300 fs and 50 𝜇J respectively. The Gaussian laser beam was circularly polarized, and, once focused onto each sample surface through a flatfield telecentric lens (Still optics), achieved a beam waist of ~15 𝜇m at the focal point. For each target, a specific matrix of laser spots (10x10 with 100 𝜇m spacing) was applied such that the effect of both the number of successive pulses and power could be studied afterwards with confidence (5 repeats of each #pulse/power combination). The general, post-laser sequence of analysis of each crater was as follows: a) optical microscopy (Keyence VHX-7000), b) White Light Interferometry (WLI, Bruker Contour GT), c) hyperspectral imaging (not reported on here), d) Scanning Electron Microscopy (SEM, Hitachi SU 5000 at 5 keV), and e) Transmission Electron Microscopy (TEM, TECNAI G2-30 Twin and FEI TITAN Themis 300). All of the experimental and characterization tools used here are located at the Université de Lille, France. Results and discussion: Using topographic data from WLI, the first implementation of an ad-hoc code (called TopoCrater) was used to generate crater topography profiles (Fig. 1b), calculate crater size metrics such as diameter:depth, and to estimate crater volume (related to quantity of ejected material). These analyses were then used to construct trends in both the number of laser pulses and laser power on each sample. 3D visual-reconstructions and topographical colourmaps (Fig. 1c) were also reproduced from WLI data, giving insight into the radial and angular morphology of each crater, and affording 'ground-truth' to morphology quantifications. As shown in Fig. 1a/c, WLI was able to resolve the complex surface structure of craters on serpentine, and identify them against other craters like those produced in olivine (which are smoother).

The SEM of laser craters also helped identify morphological trends with changing laser parameters. To probe topography inside a crater, the use of sample tilting in secondary electron (SE) mode has been particularly helpful (Fig. 1a). With SE-SEM, trends in the radial distribution of complex surface features such as melt, vesicles, bubbles, and ejecta have been studied.

By employing TEM, we have been able to spatially resolve material features and traits of interest across particular craters on each sample. This includes EDS to quantify the chemical state (not reported on here), Selective Area Electron Diffraction (SAED) to understand the crystallinity throughout a grain (Fig. 2c), and bright-and dark-field imaging to record features such as potential dislocations, fractures, or the complicated structures mentioned above (Fig. 2b). For example, after laser-cratering the olivine, Fig. 2a shows the reproduction of commonly reported laser-induced and natural Space Weathering (SpWe) features, such as an amorphous surface layer and potential nano-phase iron (shown in Fig. 2a offset from the surface, but whose spatial distribution varies radially) [4,5].

Concerning the returned samples from Ryugu, our empirical results may help interpret apparent SpWe features that have been recently discovered on the surface of some Ryugu grains [6]. In fact, surface features of our serpentine samples (Fig. 2b) seem to resemble the bubble and melt continuum structure reported in the Ryugu samples [6]. One particular note, shown in Fig. 2b, is the presence of this melt/bubble layer on presumably laser-shaded grain surfaces, despite the centre of the grain appearing unaltered. We propose this might result from the inter-granular porosity of the pressed serpentine powder target, as this is not observed in our continuous phyllosilicate chips. Understanding the crystallinity of such targets spatially, with prudent SAED (Fig. 2c), may help elucidate this phenomena.

Figure 1 .

 1 Figure 1. An example of the techniques used to characterize the topography and morphology of a crater (serpentine chip): a) an SE-SEM (tilted) image, b) a WLI-generated line-scan profile, and c) a WLIgenerated topography colour-map.

Figure 2 .

 2 Figure 2. An example of part of the approach to study crater cross-sections with TEM: a) a bright-field micrograph at the crater surface in San Carlos olivine (am = amorphous, npFe = nano-phase iron, Ol = olivine), b) a bright-field micrograph of a grain of serpentine powder at the crater surface (bb = bubble, mt = melt, Srp = serpentine), and c) an SAED pattern near the crater surface of the serpentine grain in b).

2568.pdf 53rd Lunar and Planetary Science Conference (2022)

Acknowledgments: The authors thank the I-SITE-Université de Lille Nord Europe, the Métropole Européenne de Lille, and the National Research Agency for funding. Special thanks are owed to D. Troadec for performance of FIB, and to A. Fadel for his patience in training in and performing SEM imaging. References: [1] Yada T. et al. (2021) Nat. Astron. doi.org/10.1038/s41550-021-01550-6 [2] Hamilton V. E. et al. (2019) Nat. Astron., 3, 332. [3] Kitazato K. et al. (2019) Science, 364, 272-275 [4] Weber I. et al (2020) EPSL, 530, 115884 [5] Fazio, A. et al. (2018) Icarus, 299, 240-252, [6] Noguchi T. et al (2022) LPSC LIII (this conference).