Abstract
Dynamical ensembles have been introduced to study constrained stochastic processes. In the microcanonical ensemble, the value of a dynamical observable is constrained to a given value. In the canonical ensemble a bias is introduced in the process to move the mean value of this observable. The equivalence between the two ensembles means that calculations in one or the other ensemble lead to the same result. In this paper, we study the physical conditions associated with ensemble equivalence and the consequences of non-equivalence. For continuous time Markov jump processes, we show that ergodicity guarantees ensemble equivalence. For non-ergodic systems or systems with emergent ergodicity breaking, we adapt a method developed for equilibrium ensembles to compute asymptotic probabilities while caring about the initial condition. We illustrate our results on the infinite range Ising model by characterizing the fluctuations of magnetization and activity. We discuss the emergence of non-ergodicity by showing that the initial condition can only be forgotten after a time that scales exponentially with the number of spins.
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Notes
The alternative viewpoint is to consider the non-differentiability of the CGF as the criterion for determining ensemble equivalence (merging together partial equivalence and non-equivalence).
This property is stated in problem 5, page 174 of Ref. [30].
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Acknowledgements
We acknowledge Massimiliano Esposito for his advice on this project that started when GV was a post-doctoral fellow in his research team. We thank V. Lecomte and A. Lazarescu for the insightful discussions in connection with this work.
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Appendices
Appendix A: Derivation of the Propagator for Generating Function
We look for an expression of the propagator of the generating function \(G(x_{\mathrm {f}},x_{\mathrm {i}},\kappa ,\gamma )= \big \langle e^{NT(\kappa {m}+\gamma a)}\big \rangle _{x_{\mathrm {f}},x_{\mathrm {i}}}\) as a path integral over all trajectories for the system state n. We use the discrete time \(t_k=k \mathrm {d}t\) corresponding to the k’s time step of duration \(\mathrm {d}t\), with the final time \(T=t_L = L \mathrm {d}t \). We write the system state \(n_{t_k}=n_k\) for short. This state at time \(t_k\) changes by \(n_{k}-n_{k-1}=2\epsilon _k = -2, 0\) or 2 when a spin jumps from up to down, doesn’t jump or jumps from down to up. The sum over all paths for a given initial condition is
Since jump probabilities depend on the magnetization, the path integration should include a sum on the system state \(n_k\) for all k
with the condition that one spin at most jumps at each time step. In the above sums, the index n goes from \(-N\) to N. All in all, the propagator of the generating function writes
with \(p(n_{k}|n_{k-1}) \) the probability of the transition \( n_{k-1} \rightarrow n_{k}\). Let’s use the Laplace representation of the Dirac distribution
for every integer k between 1 and L to get
In the next step, we sum over \([\epsilon _k]\) to obtain
The propagator for staying in the same state during a time step is \(p(n_{k}|n_{k}) = 1- \left( \sum \limits _{\epsilon =\pm 1}K(n_k+2\epsilon ,n_k)\right) \mathrm {d}t\), and when changing of state writes \(p(n_{k}+\epsilon |n_{k}) = K(n_k+2\epsilon ,n_k) \mathrm {d}t\). Exponentiating the product yields
The propagator of the generating function becomes in the continuous limit
where \(\int \mathcal {D} [x]\) (resp. \(\int \mathcal {D} [\mathfrak {p}]\)) is a short notation for the integral on the paths of density (resp. momentum) with given initial and final values. Finally, the path integral expression of the generating function is given by
with the action \(\mathcal {A}_{\kappa ,\gamma }\) being a functional of the paths \([x,\mathfrak {p}]\)
We have introduced the Hamiltonian function
This function is the representation in the continuous limit of a tridiagonal Metzler matrix, whose spectrum is real.Footnote 2 As a consequence, \(i \mathfrak {p}\) is a real number.
Appendix B: Wentzel–Kramers–Brillouin Approximation
We now use a saddle point integration to obtain the leading order in N of the propagator of the generating function \(G(x_{\mathrm {f}},x_{\mathrm {i}},\kappa ,\gamma )\) which is asymptotically exact in the large size limit. For each integration, we deform the contour so that it goes through the saddle point of \(\mathcal {A}_{\kappa ,\gamma } [x,\mathfrak {p},t]\). Therefore, the saddle point is the minimum of \(\mathcal {A}_{\kappa ,\gamma } [x,\mathfrak {p},t]\) solving
where \( \mathchoice{\dfrac{\delta }{\delta x_t}}{\dfrac{\delta }{\delta x_t}}{\frac{\delta }{\delta x_t}}{\frac{\delta }{\delta x_t}} \) denotes a functional derivative.
In this case, the saddle point calculation is equivalent to the WKB approximation of quantum mechanics. The propagator of the generating function will be given by the paths maximizing the action, i.e. the classical paths starting at \(x_{\mathrm {i}}\) and ending at \(x_{\mathrm {f}}\). These paths solve the Hamilton’s equations
The propagator of the generating function can be written as
where \((x,\mathfrak {p})\) correspond to the solutions of Eqs. (98)–(99) with initial and final conditions \(x_{\mathrm {i}}\) and \(x_{\mathrm {f}}\). Moreover the Hamiltonian (95) is time independent, and is hence a conserved quantity along the trajectory. Maupertuis’ action \(\int _{0}^{T} \mathrm {d}t i \mathfrak {p}_t \dot{x}_{t}\) is a function of the Hamiltonian and writes
where \(\mathfrak {p}(x)\) can be obtained by inverting Eq. (95). We observe that Maupertuis’ action is always negative in view of the clockwise motion along the orbits that can be justified by inspection of Eq. (98).
In the large size limit, we can also use a saddle point integration on \(\mathfrak {p}\) to switch to the Lagrangian framework. To do so, we look for the solution of Eq. (98). We derive a quadratic equation either for \(e^{i\mathfrak {p}_{s}}\) or for \(e^{-i\mathfrak {p}_{s}}\) whose solutions are
with \(\varphi (x,\gamma )\) defined in Eq. (59). Inserting this into Eq. (93) leads to the propagator of the generating function of Eq. (57) with Lagrangian (58).
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Vroylandt, H., Verley, G. Non-equivalence of Dynamical Ensembles and Emergent Non-ergodicity. J Stat Phys 174, 404–432 (2019). https://doi.org/10.1007/s10955-018-2186-7
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DOI: https://doi.org/10.1007/s10955-018-2186-7