Proof of the Absence of Long-Range Temporal Orders in Gibbs States

We address the question whether time translation symmetry can be spontaneously broken in a quantummany-body system. One way of detecting such a symmetry breaking is to examine the time-dependence of a correlation function. If the large-distance behavior of the correlation function exhibits a nontrivial time-dependence in the thermodynamic limit, the systemwould develop a temporal long-range order, realizing a time crystal. In an earlier publication, we sketched a proof for the absence of such time dependence in the thermal equilibriumdescribed by the Gibbs state (Watanabe and Oshikawa in Phys Rev Lett 114:251603, 2015). Here we present a complete proof and extend the argument to a more general class of stationary states than the Gibbs states.


Introduction
Time crystals are a newly proposed state of matter that spontaneously breaks the time translation symmetry. The idea of time crystals in the case of the continuous time translation symmetry was first proposed by Wilczek [1], although the validity of the concrete model in this original proposal was soon questioned in Ref. [2]. Then a no-go theorem for a wider but still restricted class of models was presented in Ref. [3]. In a more general setting, the absence of time crystalline orders in the ground state or in the Gibbs state was proven in Ref. [4] without specifying the Hamiltonian but assuming only its locality. These developments triggered further investigation of so-called Floquet time crystals or discrete time Communicated by Hal Tasaki. crystals in nonequilibrium setting [5][6][7][8][9][10] that break a discrete time translation symmetry into its subgroup. See Refs. [11][12][13] for recent reviews on this topic.
The argument for the no-go theorem at finite temperatures in Ref. [4] was based on the Lieb-Robinson bound [14,15], which was used to constrain finite time behavior of the correlation function. However, in Ref. [4], Fourier transformation of the correlation function was performed with respect to an infinitely long time, out of the validity of the constraint. This issue was recently pointed out by Ref. [13]. In this work, we present a complete version of the proof without such an issue. Furthermore, we examine the conditions on the density operator to which our argument can be straightforwardly extended. Clarifying these subtleties and settling down the limitations on what the Gibbs state and similar type of stationary states can do should in turn accelerate our exploration of new states that exhibit nontrivial temporal orders.

Setup and Statement
Let us consider a static HamiltonianĤ defined on a d-dimensional lattice that is a finite subset of Z d . We assume that the HamiltonianĤ is written as a sum of local bounded More precisely, we assume that the support of the local Hamiltonianĥ x is limited to a finite range R h from x ∈ and that the operator norm 1 ofĥ x is bounded by a constant N h . Both R h and N h are independent of the position x ∈ or the system size | |. This setting includes a wide variety of quantum spin systems, fermion systems, and "hard-core" boson systems 2 . Similarly, we consider observables (not necessarily Hermitian)Â andB written as a sum of local observables:Â The support ofâ x andb x (x ∈ ) are within a finite range R a , R b from x and their operator norm is bounded by constants N a , N b , respectively. All of these constants are independent of x or | |. We introduce the time evolution of operators for t ∈ R bŷ Our interest is in the time-dependence of the correlation function Hereρ is the Gibbs stateρ at the inverse temperature β and Z := Tr e −βĤ is the partition function. Our claim is that Â (t)B is independent of t in the thermodynamic limit | | → +∞ [4], i.e., 2.2 Proof forˇ> 0

Outline
To prove Eq. (6), it is sufficient to treat the special caseB =Â † : This is because Â (t)B can be rewritten as Once Eq. (7) is established, it applies to all four correlation functions in the right-hand side and we obtain Eq. (6). We denote by | n the eigenstate of the HamiltonianĤ with the eigenvalue E n (n ∈ N). Using the complete system and writing we get We split the summation over m and n into four intervals of E n − E m : where ε is a small positive number and K is a large positive number. Then the time-dependence of Â (t)Â † can be bounded as In the following, we derive an upper bound for each term in the right hand side one by one. The first two terms will be bounded using the Lieb-Robinson bound and the monotonically decreasing nature of the Boltzmann factor (9). The third term will be evaluated by making use of the large energy difference. Finally, the last term is trivially small because of the time-dependent factor with small energy difference. Plugging these results [Eqs. (30), (33), (38), and (39) below] into the right-hand side of Eq. (12), we get where C is a positive constant independent of the system size. Since we can take ε to be small and K to be large by choosing a sufficiently large system size | |, we obtain the desired result.

The Range (i): 2" ≤ E n − E m ≤ K
Let us start with the contribution from the range 2ε ≤ E n − E m ≤ K . To this end, we introduce a cutoff function η + ∈ C ∞ 0 (R) (i.e., an infinitely differentiable function with a compact support ) that satisfies the following conditions: The Fourier transform of η + (ω) is given bỹ which decays faster than any power of t. This can be shown by performing an integration by parts repeatedly:η which implies, for any integer ∈ N, that We consider a correlation function On one hand, we have In passing to the third line, we used ρ(E n ) < ρ(E m ) when E n > E m and the conditions (14) of η + (ω). In the last line, we defined For the Gibbs state (9), we have On the other hand, we can decompose the integral into two parts as where T is a large positive number. For the first integral in the right-hand side, we use the property Eq. (17) of the functionη + (t) as well as the trivial bound |g(t)| ≤ 2N 2 a . 4 For a given functionη + (t) with the parameters ε and K , we can find a large T such that For the second integral, we can use the Lieb-Robinson bound [14,15], from which we have [4] for system-size-independent constants C 1 and C 2 . Thus where in the second inequality we used Therefore, for any given large K and T , there exists a large volume | | such that Combining Eqs. (25) and (29) with the bound (20), we get

The Range (ii): −K ≤ E n − E m ≤ −2"
Similarly, to estimate the contribution from the range −K ≤ E n − E m ≤ −2ε, we introduce a cutoff function η − ∈ C ∞ 0 (R) that satisfies the following conditions: Here, is defined in Eq. (21). In the same way as before, we find

The Range (iii): K < |E n − E m |
The third contribution can be easily bounded by using a trick.
Thanks to the assumed locality of the Hamiltonian, this operator norm can be bounded as Note that v(R) does not grow with | |. Therefore,

The Range (iv): |E n − E m | < "
Finally, using the fact that |e i x − 1| = 2| sin x 2 | ≤ |x| for any real number x, we get This completes the verification of Eq. (13) and hence the proof of Eq. (7).

Proof forˇ= 0
Interestingly, the proof in the previous section does not apply to the infinite temperature (β = 0) where ρ(E n ) in Eq. (9) becomes constant: Here D is the dimension of the entire Hilbert space. In this special case, however, we can directly prove Eq. (6) using the clustering property of the infinite-temperature state. 5 Thus the "absence of the time crystals" also holds at the infinite temperature, consistently with the intuition that the infinite temperature is the most disordered limit. At β = 0, the equal-time correlation function trivially exhibits the locality if the support ofâ x andb y do not overlap. This implies the clustering property and the absence of any spatial long-range order. However, quantum dynamics is nontrivial even at the infinite temperature (see, for example, Ref. [16] and references therein) and the question of the time crystal is not totally trivial.
where v(R) is defined in Eq. (37). Since N a is independent of the system size, we can find | | such that 2N 2 a v(R) | | < 1 2 ε. Therefore, for a sufficiently large system size, we have δÂ δÂ † < ε.
This completes the proof of Eq. (47).