# Macroscopic Irreversibility and Decay to Kinetic Equilibrium of the 1-Body PDF for Finite Hard-Sphere Systems

Authors: Massimo Tessarotto; Claudio Cremaschini
Journal: Advances in Mathematical Physics
 (2018)

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## Abstract

The conditions for the occurrence of the so-called macroscopic irreversibility property and the related phenomenon of decay to kinetic equilibrium which may characterize the 1-body probability density function (PDF) associated with hard-sphere systems are investigated. The problem is set in the framework of the axiomatic “ab initio” theory of classical statistical mechanics developed recently and the related establishment of an exact kinetic equation realized by the Master equation for the same kinetic PDF. As shown in the paper the task involves the introduction of a suitable functional of the 1-body PDF, identified here with theMaster kinetic information. It is then proved that, provided the same PDF is prescribed in terms of suitably smooth, i.e., stochastic, solution of the Master kinetic equation, the two properties indicated above are indeed realized.


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## Body

## 1. Introduction

The axiomatic theory of classical statistical mechanics (CSM) recently proposed in a series of papers (see [1–4]) and referred to asab initio theory of CSM provides a self-consistent pathway to the kinetic theory of hard-sphere systems, as well as in principle also point particles subject to finite-range interactions [5]. Its theoretical basis and conditions of validity are indeed founded on a unique physical realization of the axioms which are set at the foundations of CSM [1–3], a fact which permits the treatment of phase-space and kinetic probability density functions (PDF) which are realized by either stochastic (i.e., ordinary) functions or distributions such as the N-body Dirac delta (or certainty function [6]). This feature is physically based being due to the prescription of the collision boundary conditions (CBC, [2]), i.e., the relationship occurring at collision events between incoming and outgoing multibody probability density functions PDF. The choice of the appropriate CBC indicated in [2], denoted as modified collision boundary condition (MCBC), is actually of crucial importance and departs from the customary realization/interpretation (of the same axioms) originally adopted in Boltzmann [7], Enskog [8, 9], and Grad [10] kinetic approaches (for a review of Grad’s kinetic theory based on CSM see also Cercignani [11, 12]). The same choice implies, in fact, a number of theoretical and physically relevant consequences. In particular, it follows that the new theory(i)
unlike Enskog theory [3] applies also tofiniteN-body hard-sphere systemsSN, namely, systems formed by N like smooth hard-spheres of diameter σ and mass m, in which the parameters (N, σ,m) remain all constant and finite [3]. On the other hand, the same particles are assumed as usual: (A) subject to instantaneous (unary, binary, and multiple) elastic collisions which leave unchanged the particles angular momenta and (B) immersed in a stationary bounded domain Ω of the Euclidean space R3 with finite canonical measure;(ii)
has led to the discovery [3] of an exact kinetic equation holding globally in time [13] (i.e., for all t∈I≡R) for these systems and denoted as Master kinetic equation (recalled in Appendix A). In other words the Master equation is nonasymptotic in character with respect to the parameters the (finite) parameters (N, σ,m). In addition the same equation holds under suitable maximal entropy conditions for the statistical treatment of the so-called Boltzmann-Sinai classical dynamical system (CDS), which implies that initial (binary or multibody) phase-space statistical correlations are assumed identically vanishing, while at the same time only suitable uniquely prescribed configuration-space correlations can arise. As such the equation generalizes and extends the validity of the Boltzmann and Enskog kinetic equations and notably applies to arbitrary 1-body PDFs which can be realized either in terms of stochastic functions or distributions;(iii)
is time-reversal invariant [4], namely, the Master kinetic equation is time-reversal (TR-) symmetric. In other words, the same equation is invariant with respect to the TR-transformation(1)τ≡t-to→τ′≡t′-to=-t+to≡-τ,r1→r1′=r1,v1→v1′=-v1.Thus, representing the absolute time ast=τ+to, with to being a prescribed (arbitrary) initial time, it follows that the TR-transformation leaves invariant the initial time to and the instantaneous position r1=r1(t≡to+τ) of an arbitrary particle, while reversing the signature (i.e., versus) of its velocity v1=v1(t≡to+τ). Accordingly, thanks to TR-symmetry the two initial-value problems associated with the Master kinetic equation in the two cases are related in such a way that, respectively, the initial 1-body PDF at time to, ρ1(N)(x1,to)≡ρ1(o)(N)(r1,v1), and the corresponding time-evolved PDF ρ1(N)(x1,t) are carried into the TR-transformed 1-body PDFs ρ1(o)TR(N)(r1′,v1′) and ρ1TR(N)(r1′,v1′,t′), respectively, prescribed according to the law(2)ρ1oNr1,v1ρ1Nx1,t≡ρ1Nr1,v1,to+τ↓ρ1oTRNr1′,v1′≡ρ1oNr1,-v1ρ1TRNr1′,v1′,t′≡ρ1Nr1,-v1,to-τ;(iv)
conserves the corresponding Boltzmann-Shannon (BS) statistical entropy [4]. This is identified with the phase-space moment(3)Sρ1Nt-∫Γ1dx1ρ1Nx1,tlnρ1Nx1,tA1,withρ1(N)(t)≡ρ1(N)(x1,t) being an arbitrary stochastic PDF solution of the Master kinetic equation and A1 an arbitrary positive constant such that the initial PDF (4)ρ1Nto≡ρ1Nx1,to=ρ10Nx1is such that the corresponding BS functionalS(ρ1(N)(to)) is defined. As a consequence it follows that an arbitrary smooth solution of the Master kinetic equation satisfies the constant H-theorem(5)∂∂tSρ1Nt=0for allt∈I≡R (see again related discussion in [4]).Based on the ab initio theory of CSM, in this paper the problem is posed of the existence of two phenomena which are expected to characterize the statistical description of finiteN-body hard-sphere systems and therefore should lay at the very foundation of CSM and kinetic theory. These are related to the physical conditions for the possible occurrence of the so-calledproperty of macroscopic irreversibility(PMI) and the consequent one represented by thedecay to kinetic equilibrium (DKE) which characterize the 1-body (kinetic) PDF in these N-body systems, i.e., when 1-body-factorized initial conditions are considered for the N-body Liouville equation [3]. The conjecture is that—in some sense in analogy with the ubiquitous character of the ergodicity property which characterizes hard-sphere systems and hence the SN-CDS [14, 15]—the occurrence of such phenomena should be independent of the number N of constituent particles of the system and therefore apply to actual physical systems for which the parameters (N,σ,m) are obviously all finite.


### 1.1. Motivations and Background

Both properties indicated above concern the statistical behavior of an ensembleSN of like particles which are advanced in time by a suitable N-body classical dynamical system, identified here with the SN-CDS. Specifically they arise in the context of the kinetic description of the same CDS, i.e., in terms of the corresponding 1-body (kinetic) probability density function (PDF) ρ1(N)(t)≡ρ1(N)(x1,t). The latter is required to belong to the functional class ρ1(N)(x1,t) of suitably smooth and strictly positive ordinary functions which are particular solutions of the relevant kinetic equation.In fact, PMI should be realized by means of a suitable, but still possibly nonunique, functional which should be globally defined in the future (i.e., for all timest≥to, being to the initial time) bounded and nonnegative, and therefore to be identified with the notion of information measure. Most importantly, however, the same functional, to be referred to here asMaster kinetic information(MKI), should also exhibit a continuously differentiable and monotonic, i.e., in particular decreasing, time-dependence.Regarding, instead, the second property of DKE this concerns the asymptotic behavior of the1-body PDF ρ1(N)(t)≡ρ1(N)(x1,t) which, accordingly, should be globally defined and decay for t→+∞ to a stationary and spatially uniform Maxwellian PDF(6)ρ1MNv1=noπ3/22To/m3/2exp⁡-mv1-Vo22To,where no>0,To>0,Vo are constant fluid fields.Both PMI and DKE correspond to physical phenomena which are actually expected to arise in disparate classicalN-body systems. The clue for their realization is represented by the ubiquitous occurrence of kinetic equilibria and consequently, in principle, also of the corresponding possible manifestation of macroscopic irreversibility and decay processes. Examples of the former ones are in principle easy to be found, ranging from neutral fluids [16] to collisional/collisionless and nonrelativistic/relativistic gases and plasmas [17–19].However, the most notable example is perhaps provided by dilute hard-sphere systems (“gases”) characterized by a large number of particles (N≡1/ε≫1) and a small (i.e., infinitesimal) diameter σ~O(ε1/2) of the same hard-spheres, for which the Boltzmann equation applies. Indeed the Boltzmann equation is actually specialized to the treatment of dilute hard-sphere systems in the Boltzmann-Grad limit discussed in the Lanford theorem [20–22] (for a detailed discussion of the topics in the context of the ab initio theory see also [4]). In such a case the 1-body PDF can be formally obtained by introducing the Boltzmann-Grad limit operator [4] (7)LBG≡limN≡1/ε→∞σ~Oε1/2,whereby the limit function ρ1(x1,t) is denoted(8)ρ1x1,t=LBGρ1Nx1,t,and ρ1(x1,t) identifies a particular solution of the Boltzmann kinetic equation.Historically, the property of irreversibility indicated above is known to be related to the Carnot’s second Law of Classical Thermodynamics. More precisely, it is related to the first-principle-proof originally attempted by Ludwig Boltzmann in 1872 [7]. Actually it is generally agreed that both phenomena lie at the very heart of Boltzmann and Grad kinetic theories [7, 10] and the related original construction of the Boltzmann kinetic equation (1872). In particular, the goal set by Boltzmann himself in his 1872 paper was the proof of Carnot’s Law providing at the same time also a possible identification of thermodynamic entropy. This was achieved in terms of what is nowadays known as Boltzmann-Shannon (BS) statistical entropy, which is identified with the phase-space moment(9)MXEρ1t≡∫Γ1dx1ρ1x1,tXEx1,t=-∫Γ1dx1ρ1x1,tlnρ1x1,tA1≡Sρ1t.Here XE(x1,t)≡-lnρ1(x1,t)/A1, ρ1(x1,t), and A1 denote, respectively, the BS entropy density, an arbitrary particular solution of the Boltzmann equation for which the same phase-space integral exists and an arbitrary positive constant. In fact, according to the Boltzmann H-theorem [7] the same functional should satisfy theentropic inequality(10)∂∂tSρ1t≥0,while, furthermore, theentropic equality condition(11)∂∂tSρ1t=0⇔ρ1x1,t=ρ1MNv1should hold. The latter equation implies therefore that, provided ρ1(t) and S(ρ1(t)) exist globally [23], then necessarily limt→+∞⁡ρ1(x1,t)=ρ1M(v1), with ρ1M(v1) denoting the stationary and spatially uniform Maxwellian PDF (6).In this reference, however, the question arises of the precise characterization of the concept of irreversibility, i.e., whether it should be regarded as a purely macroscopic phenomenon (“macroscopic irreversibility”), i.e., affecting only the BS entropyS(ρ1(t)) through the Boltzmann H-theorem indicated above, or microscopic in the sense that the same Boltzmann equation should be considered as irreversible (“microscopic irreversibility”). Thus, in principle, in the second case the further issue emerges of the possible physical origin of microscopic irreversibility in special reference to the Lanford’s derivation of the Boltzmann equation and subsequent related comments discussed respectively by Uffink and Valente and Ardourel in [24, 25] (see also Drory [26]).However, as shown in [4], the Boltzmann equation is actually TR-symmetric. Such a conclusion is of basic importance since it overcomes the so-called Loschmidt paradox, i.e., the objection raised by Loschmidt in 1876 [27] regarding the original Boltzmann formulation of his namesake kinetic equation and H-theorem [7]. In fact, Loschmidt claimed that the Boltzmann H-theorem inequality should change sign under time-reversal and thus violate the microscopic time-reversibility of the underlying hard-sphere classical dynamical system. In his long-pondered reply given in 1896 [28] Boltzmann himself introduced what was later referred to as the modified form of the Boltzmann H-theorem [29].The key implication is therefore that, in contrast to Boltzmann’s own statement and the traditional subsequent mainstream literature interpretation (see, for example, by Cercignani, Lebowitz in [30, 31] and more recently the review given by Gallavotti [32]), the Boltzmann H-theorem indicated—together with the modified form indicated above—cannot be interpreted as an intrinsic irreversibility property occurring at the microscopic level, namely, holding for the Boltzmann equation itself. On the contrary, consistent with the physical interpretation of the Loschmidt paradox provided in [4], this must be regarded only asproperty of macroscopic irreversibility (or PMI) of the 1-body PDF solution of the Boltzmann equation. In other words, the Boltzmann inequality (10) necessarily holdsindependent of the orientation of the time axis (arrow of time) and therefore cannot represent a true (i.e., microscopic) property which as such should uniquely determine the arrow of time.Nevertheless, the possible realization of either PMI or DKE is more subtle. In fact they actually depend in a critical way on the prescription of the functional classρ(N)(x1,t), so that their occurrence is actually nonmandatory. Indeed, both cannot occur—in principle also for Boltzmann and Grad kinetic theories—if the N-body probability density function ρ(N)(x,t) is identified with the deterministic N-body PDF [1], namely, the N-body phase-space Dirac delta. This is defined as δ(x-x(t))≡∏1=1,Nδ(xi-xi(t)), with x≡x1,…,xN denoting the state of the N-body system and x(t)≡x1(t),…,xN(t) is the image of an arbitrary initial state x(to)≡xo generated by the same N-body CDS. That such a PDF necessarily must realize an admissible particular solution of the N-body Liouville equation follows, in fact, as a straightforward consequence of the axioms of classical statistical mechanics [1].Despite these premises, however, the case of a finite Boltzmann-Sinai CDS, which is characterized by a finite number of particlesN and/or a finite-size of the hard-spheres and/or a dense or locally dense system, is more subtle and—as explained below—even unprecedented since it has actually remained unsolved to date. The reasons are as follows. First, Boltzmann and Grad kinetic theories are inapplicable to the finite Boltzmann-Sinai CDS. Second, the Boltzmann-Shannon entropy associated with an arbitrary particular solution ρ(N)(t)≡ρ(N)(x1,t) of the Master kinetic equation, i.e., the functional S(ρ1(N)(t))≡MXE(ρ1(N)(t)), in contrast to S(ρ1(t))≡MXE(ρ1(t)), is exactly conserved in the sense that identically(12)∂Sρ1Nt∂t≡0must hold. As a consequence the validity itself of Boltzmann H-theorem breaks down in the case of the Master kinetic equation. Third, an additional motivation is provided by the conjecture that both PMI and DKE might occur only if the Boltzmann-Grad limit is actually performed, i.e., only in validity of Boltzmann equation and H-theorem.Hence the question which arises is whether in the case of a finite Boltzmann-Sinai CDS the phenomenon of DKE may still arise. Strong indications seem to be hinting at such a possibility. In this regard the example-case which refers to the statistical description of a Navier-Stokes fluid described by the incompressible Navier-Stokes equations (INSE) in terms of the Master kinetic equation is relevant and suggests that this may be indeed the case. In fact, thanks also to comparisons with the mean-field inverse kinetic approach to INSE [16], in such a case the decay of the fluid velocity field occurring in a bounded domain necessarily demands the existence of DKE. In other words, in the limit t→+∞ the 1-body PDF must decay uniformly to the stationary and spatially uniform Maxwellian PDF (6).However, besides the construction of the kinetic equation appropriate for such a case, a further unsolved issue lies in the determination of the functional classρ1(N)(x1,t) for which both PMI and DKE should/might be realized. In particular, the possible occurrence of both PMI and DKE should correspond to suitably smooth, but nonetheless still arbitrary, initial conditions ρ1(N)(x1,to). These should warrant that in the limit t→+∞, ρ1(N)(x1,t) uniformly converges to the spatially homogeneous and stationary Maxwellian PDF ρ1M(v1) (6). Such a result, however, is highly nontrivial since it should rely on the establishment of a global existence theorem for the same 1-body PDF ρ1(N)(x1,t)—namely, holding in the whole time axis I≡R, besides the same 1-body phase space Γ1—for the involved kinetic equation which is associated with the SN-CDS. In the context of the Boltzmann equation in particular, despite almost-endless efforts this task has actually not been accomplished yet, the obstacle being intrinsically related to the asymptotic nature of the Boltzmann equation [13]. In fact for the same equation it is not known in satisfactory generality whether smooth enough solutions of the same equation exist which satisfy the H-theorem inequality and decay asymptotically to kinetic equilibrium [23, 30].


### 1.2. Goals and Organization of the Paper

Based on these premises, the crucial new results that we intend to display in this paper concern theproof-of-principle of two phenomena which are expected to characterize the statistical description of finite N-body hard-sphere systems and therefore should lay at the very foundation of classical statistical mechanics and kinetic theory alike. These are related to the physical conditions for the possible occurrence of both PMI and the consequent one represented by the possible occurrence of DKE which should characterize the kinetic PDF in these systems. These phenomena are well known to occur in the case of dilute hard-sphere systems, i.e., in the Boltzmann-Grad limit. In particular, for an exhaustive treatment of the related issues which arise in the context of the ab initio theory we refer to discussions reported in [4]. Nevertheless, as indicated above, their existence in the case of finite hard-sphere systems is partly motivated by a previous investigation dealing with the kinetic description incompressible Navier-Stokes granular fluids [33].Therefore, main goal of the paper is to show that these properties actually emerge as necessary implications of the ab initio theory of CSM. Incidentally, in doing so, the Master kinetic equation must be necessarily adopted. In fact, the finiteness requirement on theSN-CDS rules out for further possible consideration either the Boltzmann or the Enskog kinetic equations, these equations being inapplicable to the treatment of systems of this type [3]. Specifically, in the following the case N>2 is considered everywhere, which is by far the most physically relevant one. In this occurrence, in fact, nontrivial 2-body occupation coefficients arise (see related notations which are applicable for N>2 recalled in Appendices A and B below). For completeness the case N=2 is nevertheless briefly discussed in Appendix D.For this purpose, first, in Section2, the MKI functional is explicitly determined. We display in particular its construction method (seeNo.#1- #4 MKI Prescriptions). Based on the theory of the Master kinetic equation earlier developed [3] and suitable integral and differential identities (see Appendices A, B, and C), the properties of the MKI functional are investigated. These concern in particular the establishment of appropriate inequalities holding for the same functional (Theorem 1, Section 2.1), the signature of the time derivative of the same functional (Theorem 2, Section 2.2) and the property of DKE holding for a suitable class of 1-body PDFs (Theorem 3, Section 2.3). In the subsequent Sections 3 and 4, the issue of the consistency of the phenomena of PMI and DKE with microscopic dynamics is posed together with the physical interpretation and implications of the theory. The goal is to investigate the relationship of the DKE theory developed here with the microscopic reversibility principle and the Poincaré recurrence theorem. Finally in Section 5 the conclusions of the paper are drawn and possible applications/developments of the theory are pointed out.


## 1.1. Motivations and Background

Both properties indicated above concern the statistical behavior of an ensembleSN of like particles which are advanced in time by a suitable N-body classical dynamical system, identified here with the SN-CDS. Specifically they arise in the context of the kinetic description of the same CDS, i.e., in terms of the corresponding 1-body (kinetic) probability density function (PDF) ρ1(N)(t)≡ρ1(N)(x1,t). The latter is required to belong to the functional class ρ1(N)(x1,t) of suitably smooth and strictly positive ordinary functions which are particular solutions of the relevant kinetic equation.In fact, PMI should be realized by means of a suitable, but still possibly nonunique, functional which should be globally defined in the future (i.e., for all timest≥to, being to the initial time) bounded and nonnegative, and therefore to be identified with the notion of information measure. Most importantly, however, the same functional, to be referred to here asMaster kinetic information(MKI), should also exhibit a continuously differentiable and monotonic, i.e., in particular decreasing, time-dependence.Regarding, instead, the second property of DKE this concerns the asymptotic behavior of the1-body PDF ρ1(N)(t)≡ρ1(N)(x1,t) which, accordingly, should be globally defined and decay for t→+∞ to a stationary and spatially uniform Maxwellian PDF(6)ρ1MNv1=noπ3/22To/m3/2exp⁡-mv1-Vo22To,where no>0,To>0,Vo are constant fluid fields.Both PMI and DKE correspond to physical phenomena which are actually expected to arise in disparate classicalN-body systems. The clue for their realization is represented by the ubiquitous occurrence of kinetic equilibria and consequently, in principle, also of the corresponding possible manifestation of macroscopic irreversibility and decay processes. Examples of the former ones are in principle easy to be found, ranging from neutral fluids [16] to collisional/collisionless and nonrelativistic/relativistic gases and plasmas [17–19].However, the most notable example is perhaps provided by dilute hard-sphere systems (“gases”) characterized by a large number of particles (N≡1/ε≫1) and a small (i.e., infinitesimal) diameter σ~O(ε1/2) of the same hard-spheres, for which the Boltzmann equation applies. Indeed the Boltzmann equation is actually specialized to the treatment of dilute hard-sphere systems in the Boltzmann-Grad limit discussed in the Lanford theorem [20–22] (for a detailed discussion of the topics in the context of the ab initio theory see also [4]). In such a case the 1-body PDF can be formally obtained by introducing the Boltzmann-Grad limit operator [4] (7)LBG≡limN≡1/ε→∞σ~Oε1/2,whereby the limit function ρ1(x1,t) is denoted(8)ρ1x1,t=LBGρ1Nx1,t,and ρ1(x1,t) identifies a particular solution of the Boltzmann kinetic equation.Historically, the property of irreversibility indicated above is known to be related to the Carnot’s second Law of Classical Thermodynamics. More precisely, it is related to the first-principle-proof originally attempted by Ludwig Boltzmann in 1872 [7]. Actually it is generally agreed that both phenomena lie at the very heart of Boltzmann and Grad kinetic theories [7, 10] and the related original construction of the Boltzmann kinetic equation (1872). In particular, the goal set by Boltzmann himself in his 1872 paper was the proof of Carnot’s Law providing at the same time also a possible identification of thermodynamic entropy. This was achieved in terms of what is nowadays known as Boltzmann-Shannon (BS) statistical entropy, which is identified with the phase-space moment(9)MXEρ1t≡∫Γ1dx1ρ1x1,tXEx1,t=-∫Γ1dx1ρ1x1,tlnρ1x1,tA1≡Sρ1t.Here XE(x1,t)≡-lnρ1(x1,t)/A1, ρ1(x1,t), and A1 denote, respectively, the BS entropy density, an arbitrary particular solution of the Boltzmann equation for which the same phase-space integral exists and an arbitrary positive constant. In fact, according to the Boltzmann H-theorem [7] the same functional should satisfy theentropic inequality(10)∂∂tSρ1t≥0,while, furthermore, theentropic equality condition(11)∂∂tSρ1t=0⇔ρ1x1,t=ρ1MNv1should hold. The latter equation implies therefore that, provided ρ1(t) and S(ρ1(t)) exist globally [23], then necessarily limt→+∞⁡ρ1(x1,t)=ρ1M(v1), with ρ1M(v1) denoting the stationary and spatially uniform Maxwellian PDF (6).In this reference, however, the question arises of the precise characterization of the concept of irreversibility, i.e., whether it should be regarded as a purely macroscopic phenomenon (“macroscopic irreversibility”), i.e., affecting only the BS entropyS(ρ1(t)) through the Boltzmann H-theorem indicated above, or microscopic in the sense that the same Boltzmann equation should be considered as irreversible (“microscopic irreversibility”). Thus, in principle, in the second case the further issue emerges of the possible physical origin of microscopic irreversibility in special reference to the Lanford’s derivation of the Boltzmann equation and subsequent related comments discussed respectively by Uffink and Valente and Ardourel in [24, 25] (see also Drory [26]).However, as shown in [4], the Boltzmann equation is actually TR-symmetric. Such a conclusion is of basic importance since it overcomes the so-called Loschmidt paradox, i.e., the objection raised by Loschmidt in 1876 [27] regarding the original Boltzmann formulation of his namesake kinetic equation and H-theorem [7]. In fact, Loschmidt claimed that the Boltzmann H-theorem inequality should change sign under time-reversal and thus violate the microscopic time-reversibility of the underlying hard-sphere classical dynamical system. In his long-pondered reply given in 1896 [28] Boltzmann himself introduced what was later referred to as the modified form of the Boltzmann H-theorem [29].The key implication is therefore that, in contrast to Boltzmann’s own statement and the traditional subsequent mainstream literature interpretation (see, for example, by Cercignani, Lebowitz in [30, 31] and more recently the review given by Gallavotti [32]), the Boltzmann H-theorem indicated—together with the modified form indicated above—cannot be interpreted as an intrinsic irreversibility property occurring at the microscopic level, namely, holding for the Boltzmann equation itself. On the contrary, consistent with the physical interpretation of the Loschmidt paradox provided in [4], this must be regarded only asproperty of macroscopic irreversibility (or PMI) of the 1-body PDF solution of the Boltzmann equation. In other words, the Boltzmann inequality (10) necessarily holdsindependent of the orientation of the time axis (arrow of time) and therefore cannot represent a true (i.e., microscopic) property which as such should uniquely determine the arrow of time.Nevertheless, the possible realization of either PMI or DKE is more subtle. In fact they actually depend in a critical way on the prescription of the functional classρ(N)(x1,t), so that their occurrence is actually nonmandatory. Indeed, both cannot occur—in principle also for Boltzmann and Grad kinetic theories—if the N-body probability density function ρ(N)(x,t) is identified with the deterministic N-body PDF [1], namely, the N-body phase-space Dirac delta. This is defined as δ(x-x(t))≡∏1=1,Nδ(xi-xi(t)), with x≡x1,…,xN denoting the state of the N-body system and x(t)≡x1(t),…,xN(t) is the image of an arbitrary initial state x(to)≡xo generated by the same N-body CDS. That such a PDF necessarily must realize an admissible particular solution of the N-body Liouville equation follows, in fact, as a straightforward consequence of the axioms of classical statistical mechanics [1].Despite these premises, however, the case of a finite Boltzmann-Sinai CDS, which is characterized by a finite number of particlesN and/or a finite-size of the hard-spheres and/or a dense or locally dense system, is more subtle and—as explained below—even unprecedented since it has actually remained unsolved to date. The reasons are as follows. First, Boltzmann and Grad kinetic theories are inapplicable to the finite Boltzmann-Sinai CDS. Second, the Boltzmann-Shannon entropy associated with an arbitrary particular solution ρ(N)(t)≡ρ(N)(x1,t) of the Master kinetic equation, i.e., the functional S(ρ1(N)(t))≡MXE(ρ1(N)(t)), in contrast to S(ρ1(t))≡MXE(ρ1(t)), is exactly conserved in the sense that identically(12)∂Sρ1Nt∂t≡0must hold. As a consequence the validity itself of Boltzmann H-theorem breaks down in the case of the Master kinetic equation. Third, an additional motivation is provided by the conjecture that both PMI and DKE might occur only if the Boltzmann-Grad limit is actually performed, i.e., only in validity of Boltzmann equation and H-theorem.Hence the question which arises is whether in the case of a finite Boltzmann-Sinai CDS the phenomenon of DKE may still arise. Strong indications seem to be hinting at such a possibility. In this regard the example-case which refers to the statistical description of a Navier-Stokes fluid described by the incompressible Navier-Stokes equations (INSE) in terms of the Master kinetic equation is relevant and suggests that this may be indeed the case. In fact, thanks also to comparisons with the mean-field inverse kinetic approach to INSE [16], in such a case the decay of the fluid velocity field occurring in a bounded domain necessarily demands the existence of DKE. In other words, in the limit t→+∞ the 1-body PDF must decay uniformly to the stationary and spatially uniform Maxwellian PDF (6).However, besides the construction of the kinetic equation appropriate for such a case, a further unsolved issue lies in the determination of the functional classρ1(N)(x1,t) for which both PMI and DKE should/might be realized. In particular, the possible occurrence of both PMI and DKE should correspond to suitably smooth, but nonetheless still arbitrary, initial conditions ρ1(N)(x1,to). These should warrant that in the limit t→+∞, ρ1(N)(x1,t) uniformly converges to the spatially homogeneous and stationary Maxwellian PDF ρ1M(v1) (6). Such a result, however, is highly nontrivial since it should rely on the establishment of a global existence theorem for the same 1-body PDF ρ1(N)(x1,t)—namely, holding in the whole time axis I≡R, besides the same 1-body phase space Γ1—for the involved kinetic equation which is associated with the SN-CDS. In the context of the Boltzmann equation in particular, despite almost-endless efforts this task has actually not been accomplished yet, the obstacle being intrinsically related to the asymptotic nature of the Boltzmann equation [13]. In fact for the same equation it is not known in satisfactory generality whether smooth enough solutions of the same equation exist which satisfy the H-theorem inequality and decay asymptotically to kinetic equilibrium [23, 30].


## 1.2. Goals and Organization of the Paper

Based on these premises, the crucial new results that we intend to display in this paper concern theproof-of-principle of two phenomena which are expected to characterize the statistical description of finite N-body hard-sphere systems and therefore should lay at the very foundation of classical statistical mechanics and kinetic theory alike. These are related to the physical conditions for the possible occurrence of both PMI and the consequent one represented by the possible occurrence of DKE which should characterize the kinetic PDF in these systems. These phenomena are well known to occur in the case of dilute hard-sphere systems, i.e., in the Boltzmann-Grad limit. In particular, for an exhaustive treatment of the related issues which arise in the context of the ab initio theory we refer to discussions reported in [4]. Nevertheless, as indicated above, their existence in the case of finite hard-sphere systems is partly motivated by a previous investigation dealing with the kinetic description incompressible Navier-Stokes granular fluids [33].Therefore, main goal of the paper is to show that these properties actually emerge as necessary implications of the ab initio theory of CSM. Incidentally, in doing so, the Master kinetic equation must be necessarily adopted. In fact, the finiteness requirement on theSN-CDS rules out for further possible consideration either the Boltzmann or the Enskog kinetic equations, these equations being inapplicable to the treatment of systems of this type [3]. Specifically, in the following the case N>2 is considered everywhere, which is by far the most physically relevant one. In this occurrence, in fact, nontrivial 2-body occupation coefficients arise (see related notations which are applicable for N>2 recalled in Appendices A and B below). For completeness the case N=2 is nevertheless briefly discussed in Appendix D.For this purpose, first, in Section2, the MKI functional is explicitly determined. We display in particular its construction method (seeNo.#1- #4 MKI Prescriptions). Based on the theory of the Master kinetic equation earlier developed [3] and suitable integral and differential identities (see Appendices A, B, and C), the properties of the MKI functional are investigated. These concern in particular the establishment of appropriate inequalities holding for the same functional (Theorem 1, Section 2.1), the signature of the time derivative of the same functional (Theorem 2, Section 2.2) and the property of DKE holding for a suitable class of 1-body PDFs (Theorem 3, Section 2.3). In the subsequent Sections 3 and 4, the issue of the consistency of the phenomena of PMI and DKE with microscopic dynamics is posed together with the physical interpretation and implications of the theory. The goal is to investigate the relationship of the DKE theory developed here with the microscopic reversibility principle and the Poincaré recurrence theorem. Finally in Section 5 the conclusions of the paper are drawn and possible applications/developments of the theory are pointed out.


## 2. Axiomatic Prescription of the MKI Functional

In view of the considerations given above in this section the problem is posed of the explicit realization of the MKI functional in terms of suitable axiomatic prescriptions. The same functional, denotedIM(ρ1(N)(t)), should depend on the 1-body PDF ρ1(N)(t), with ρ1(N)(t)≡ρ1(N)(x1,t) being identified with a particular solution of the Master kinetic equation (see (A.4) in Appendix A holding for N>2 and Appendix D for the case N=2).Unlike Boltzmann kinetic equation, the Master kinetic equation actually deals with the treatment of finite hard-sphereN-body systems, i.e., in which both the number of particles N and their diameter σ remain finite [3]. To achieve such a goal suitably prescribed physical collision boundary conditions (CBC) of the N-body PDF need to be adopted. More precisely, this concerns the prescription for arbitrary collision events of the relationship between incoming (−) and outgoing (+) PDFs, i.e., respectively, the left and right limits ρ(±)(N)(x(±)(ti),ti)=limt→ti(±)⁡ρ(N)(x(t),t), with x(±)(ti)=limt→ti(±)⁡x(t) denoting the corresponding incoming (−) and outgoing (+) states. In particular, upon invoking due to causality the assumption of left-continuity, i.e., the requirement(13)ρ-Nx-ti,ti≡ρNx-ti,ti,the incoming PDF is required to coincide with the same N-body PDF evaluated in terms of the incoming state and time [1, 3]. Hence, as recalled in Appendix C (see also [2]) from (C.1) it follows that the so-called causal form of the modified collision boundary condition (MCBC [2])(14)ρ+Nx+ti,ti=ρNx+ti,tiis mandatory. A further important requirement concerns precisely setting also the relatedfunctional class of admissible solutionsρ1(N)(x1,t) in such a way that, besides ρ1(N)(t), also the same functional IM(ρ1(N)(t)) exists globally for arbitrary t∈I≡R. For definiteness, we shall consider for this purpose the case of 1-body PDFs which satisfy the initial condition(15)ρ1Nto≡ρ1Nx1,to=ρ1oNx1,with ρ1(o)(N)(x1) belonging to the functional class ofstochastic1-body PDFsρ1(N)(to). For a generic t∈I belonging to the time axis I≡R this is the ensemble of 1-body PDFs ρ1(N)(t) which are, respectively, (A) smoothly differentiable; (B) strictly positive; (C) summable, in the sense that the velocity—or phase-space—moments for the same PDF exist which correspond either to arbitrary monomial functions of v1 (or its components v1i, for i=1,2,3) or to the entropy density lnρ1(N)(t), thus yielding the Boltzmann-Shannon (BS) entropy evaluated in terms of ρ1(N)(t).Concerning the choice of the settingρ1(N)(to) the following remarks are in order. As a first remark, the previous requirements (A), (B), and (C) for ρ1(N)(to), together with validity of MCBC (14), actually should warrant that the corresponding solution of the Master kinetic equation ρ1(N)(t) exists globally in the extended phase-space x1,t∈Γ1×I and that for all t∈I the same PDF belongs to the class of stochastic PDFs ρ1(N)(x1,t) indicated above and also fulfills identically the constant H-theorem (12). Indeed, one can show [13] that global existence of solutions for the Master kinetic equation follows in elementary way from the N-body Liouville equation. Indeed, an arbitrary 1-body PDF which is a particular solution of the Master kinetic equation realizes by construction also a particular factorized solution of the N-body Liouville equation, i.e., of the N-body PDF [3]. The same PDF evolves uniquely in time along arbitrary phase-space Lagrangian trajectories, its Lagrangian time evolution being determined at arbitrary collision times by MCBC (14) [13].As a second remark,the validity of assumptions (A) and (B) for ρ1(o)(N)(x1) and ρ1(N)(x1,t) implies also suitable assumptions to apply all t∈I for the local characteristic scale length Lρ which characterize the same PDF ρ1(N)(x1,t). More precisely, this is associated with the spatial variations of the 1-body PDF prescribed as(16)Lρt=infx1∈Γ1∂lnρ1Nx1,t∂r1-1,which necessarily assumed nonzero at all time t∈I. Hence ∂lnρ1(N)x1,t/∂r1 is assumed to be bounded for all x1,t spanning the extended 1-body phase space Γ1(1)×I.Finally, as a third remark (see also the further related discussion in Theorem2 below), the previous requirements are expected to warrant also the global existence of the MKI functional IM(ρ1(N)(t)), so that ρ1(N)(x1,t) effectively realizes the functional class of admissible solutions indicated above.Given these premises let us pose now the problem of the identification of the functionalIMρ1o(N)(x1), based on the introduction of ’ad hoc’ physical requirements, to be referred to here asMKI Prescriptions No.#1-#4. The prescriptions are as follows:(i)
MKI Prescription No.#1:the first one is that the functional IM(ρ1(N)(t)) should be determined in such a way that the existence of IMρ1o(N)(x1) at a suitable initial time to∈I should warrant also that IM(ρ1(N)(t)) must necessarily exist globally in the future, i.e., for all t≥to. As a consequence the functional class ρ1(N)(x1,t) must be suitably prescribed.(ii)
MKI Prescription No.#2:second, we shall require that IM(ρ1(N)(t)) is real, nonnegative, and bounded in the sense that(17)0≤IMρ1Nt≤1.This implies thatIM(ρ1(N)(t)) can be interpreted as an information measure associated with the 1-body PDF ρ1(N)(t)≡ρ1(N)(x1,t). For this reason the previous inequalities will be referred to asinformation-measure inequalities.(iii)
MKI Prescription No.#3:third, for consistency with the property of macroscopic irreversibility, IM(ρ1(N)(t)) is prescribed in terms of a smoothly time-differentiable and monotonically time-decreasing functional in the sense that in the same time-subset the inequality(18)∂∂tIMρ1Nt≤0should identically apply∀t≥to, so that(19)0≤IMρ1Nt≤IMρ1oNx1≤1.This implies thatIM(ρ1(N)(t)) is also globally defined for all t∈I≡R with t≳to. In addition, if ∂/∂tIM(ρ1(N)(t))t=to≠0, without loss of generality its initial value IMρ1o(N)(x1), can always be set such that(20)IMρ1oNx1=1.(iv)
MKI Prescription No.#4:fourth, in order to warrant the existence of DKE we shall require the functional IM(ρ1(N)(t)) to be prescribed in such a way that at an arbitrary time t∈I, with t≳to, the vanishing of both IM(ρ1(N)(t)) and its time derivative ∂/∂tIM(ρ1(N)(t)) should occur if and only if the 1-body PDF solution of the Master kinetic equation coincides with kinetic equilibrium. As a consequence, for the functional IM(ρ1(N)(t)) the following propositions should be equivalent:(21)IMρ1Nt=0∂∂tIMρ1Nt=0⇕ρ1Nx1,t≡ρ1MNv1,withρ1M(N)(v1) being a kinetic equilibrium PDF of the form (6).The immediate obvious implication of the previous prescriptions is that—provided a nontrivial realization of the MKI can be found in the functional class ρ1o(N)(x1)—the existence of both PMI and DKE for the Master kinetic equation is actually established. In the sequel the goal is to show, in particular, that the MKI functional can be identified by means of the prescription(22)IMρ1Nt,b≡KMρ1Nt,bKMo,where KM(ρ1(N)(t),b) and KMo denote, respectively, a suitable (and possibly nonunique) moment-dependent phase-space functional and an appropriate normalization constant to be chosen in such a way to satisfy all the MKI prescriptions indicated above. In particular, as shown below, an admissible choice for KM(ρ1(N)(t),b) and KMo is provided by(23)KMρ1Nt,b=-∫Γ11dx1Θ¯1∂Ωr¯1Mv1,bρ1Nx1,tρ^1Nx1,t∂2ρ^1Nx1,t∂r1·∂r1,KMo=sup⁡1,KMρ1oNx1,b,while M(v1,b) denotes thedirectional kinetic energy (along the unit vector b) carried by particle 1, namely, the dynamical variable(24)Mv1,b≡v1·b2,with b denoting a still arbitrary constant unit vector. Hence,(25)Mv1,v2,b=12Mv1,b+Mv2,bidentifies the correspondingtotal directional kinetic energy carried by particles 1 and 2. Here the remaining notation is standard. Thus, ρ1(N)(t)≡ρ1(N)(x1,t), ρ1o(N)(x1), and ρ^1(N)(t)≡ρ^1(N)(x1,t) are, respectively, the 1-body PDF solution of the initial problem associated with the Master kinetic equation (see (A.4) in Appendix A), the initial PDF, and the renormalized 1-body PDF(26)ρ^1Nx1,t≡ρ1Nx1,tk1Nr1,t,while furthermore k1(N)(r1,t) is the 1-body occupation coefficient recalled in Appendix B (see (B.1)). As a consequence in the previous equation it follows that ρ1(N)(x1,t)/ρ^1(N)(x1,t)≡k1(N)(r1,t). Furthermore, Θ¯1(∂Ω)(r¯1) is the boundary theta function given by (A.10) (see Appendix A). Finally, regarding the initial value KM(ρ1o(N)(x1),b) it follows that if, respectively, KM(ρ1o(N)(x1),b)≥1 or(27)0≤KMρ1oNx1,b<1,then correspondingly one obtains, consistent with (19), that the initial value of MKI functional IM(ρ1o(N),b) is (28)IMρ1oN,b=1,KMρ1oNx1,b.


### 2.1. Proof of the Nonnegativity of the MKI Information Measure

The strategy adopted for the proof of the MKI Prescriptions No.#1 and No.#2 is to show initially the validity of the information-measure left inequality in (17), namely, that IM(ρ1(N)(t),b) cannot acquire negative values for arbitrary t≥to. The result is established by the following theorem.Theorem 1 (nonnegativity ofKM(ρ1o(N)(x1),b), KM(ρ1(N)(t),b), and IM(ρ1(N)(t),b)).
Let us assume thatρ1(N)(x1,t) is an arbitrary stochastic and suitably smoothly differentiable, particular solution of the Master kinetic equation (A.4) prescribed so that the integral (23) expressed in terms of the initial PDF, namely, KM(ρ1o(N)(x1),b), is nonvanishing. Then, it follows necessarily that(i)
PropositionP11:(29)KMρ1oNx1,b>0.(ii)
PropositionP12: the corresponding time-evolved functional KM(ρ1(N)(t),b) for all t∈I with t>to is such that(30)KMρ1Nt,b≥0.(iii)
PropositionP13: for all t∈I with t>to the functional IM(ρ1(N)(t),b) fulfills the inequality(31)IMρ1Nt,b≥0.(iv)
PropositionP14: the following necessary and sufficient condition holds at a given time t∈I with t≥to:(32)KMρ1Nt,b=0⇔ρ1Nx1,t≡ρ1MNv1.Proof.
One first notices thatKM(ρ1(N)(t),b) can be equivalently written in the form(33)KMρ1Nt,b≡-∫Γ11dx1Θ¯1∂Ωr¯1Mv1,bk1Nr1,t∂2ρ^1Nx1,t∂r1·∂r1, where in order that the same functional exists it is obvious that the renormalized 1-body PDF ρ^1(N)(x1,t) must be of class C(2). Integrating by parts and noting that the gradient term ∂Θ¯1(∂Ω)(r¯)/∂r1 gives a vanishing contribution to the phase-space integral, this yields equivalently(34)KMρ1Nt,b≡∫Γ11dx1Θ¯1∂Ωr¯1Mv1,b∂k1Nr1,t∂r1·∂ρ^1Nx1,t∂r1. Therefore, upon invoking (B.9) reported in Appendix B, direct substitution delivers(35)KMρ1Nt,b=N-1∫Γ11dx1Θ¯1∂Ωr¯1Mv1,b∂ρ^1Nx1,t∂r1·∫Γ12dx2n12×δr2-r1-σΘ¯2∂Ωr¯2ρ^1Nx2,tk2Nr1,r2,t.Next, invoking the identity n12δr2-r1-σ=-∂/∂r2Θ¯r2-r1-σ and noting again that ∂/∂r2Θ¯2(∂Ω)(r¯) gives vanishing contribution, one can perform a further integration by parts with respect to r2. This permits to cast the rhs of previous equation in the form(36)KMρ1Nt,b≡KM1ρ1Nx1,t,b+ΔKM1ρ1Nx1,t,b.Here the two terms on the rhs of (36) are defined as follows: (1) the first term KM(1)(ρ1(N)(t),b) is symmetric and nonnegative, so that it can be expressed so to carry the total directional kinetic energy M(v1,v2,b) of particles 1 and 2 (see (25)). Hence, it takes the form(37)KM1ρ1Nx1,t,b=N-1∫Γ11dx1∫Γ12dx2Θ¯1∂Ωr¯1Θ¯2∂Ωr¯2×∂ρ^1Nx1,t∂r1·∂ρ^1Nx2,t∂r2k2Nr1,r2,tMv1,v2,bΘ¯r2-r1-σ.(2) The second term ΔKM(1)(ρ1(N)(x1,t),b) reads instead(38)ΔKM1ρ1Nx1,t,b≡N-1∫Γ11dx1Θ¯1∂Ωr¯1∂ρ^1Nx1,t∂r1·∫Γ12dx2×Θ¯2∂Ωr¯2Mv1,bΘ¯r2-r1-σρ^1Nx2,t∂∂r2k2Nr1,r2,t,where ∂/∂r2k2(N)(r1,r2,t) is given by the differential identity (B.10) reported in Appendix B. Thus, upon invoking the identity n23δr3-r2-σ=-∂/∂r3Θ¯r3-r2-σ, one notices that an integration by parts can be performed also with respect to r3. This means that a procedure analogous to the one used for the calculation of ∂k1(N)(r1,t)/∂r1 can be invoked and iterated at all orders, i.e., up to the (N-1)-body occupation coefficient (see (B.11) in Appendix B). As a consequence the functional KM(ρ1(N)(t),b) can be represented in terms of a finite sum of the form KM(ρ1(N)(t),b)≡∑j=1,N-1KM(j)(ρ1(N)(x1,t),b) in which each term of the sum KM(j)(ρ1(N)(x1,t),b) is nonnegative and symmetric. This implies therefore that the same functional KM(ρ1(N)(t),b) can be cast in the form(39)KMρ1Nt,b=∫Γ11dx1∫Γ12dx2Θ¯1∂Ωr¯1Θ¯2∂Ωr¯2Mv1,v2,b×Fr1,r2,t∂ρ^1Nx1,t∂r1·∂ρ^1Nx2,t∂r2Θ¯r2-r1-σ,with M(v1,v2,b)≥0 being the total directional kinetic energy (25) and F(r1,r2,t) a suitable real scalar kernel which is symmetric in the variables r1 and r2. Hence KM(ρ1(N)(t),b) actually defines a nonnegative functional. This proves the validity of the inequality (31) (Proposition P11).
In a similar way also the remaining propositions can be established. In fact, invoking (28) it follows that the inequalities (30) and (31)—and hence also Propositions P12 and P13—manifestly hold too. Finally, regarding the proof of Proposition P14, one notices that KM(ρ1(N)(t),b)≡0 if and only if identically ∂/∂r1ρ^1(N)(x1,t)≡0. Since ρ^1(N)(x1,t) is by construction a solution of the Master kinetic equation it follows that this requires necessarily that ρ^1(N)(x1,t) must coincide with the local Maxwellian ρ1M(N)(v1) (see (6)) and hence (32) must hold too under the same realization (Proposition P14).The conclusion is therefore that the definition of the MKI functional (22) given above in terms of KM(ρ1(N)(t),b) and KMo (see (23)) is indeed consistent with the physical prerequisites represented by the MKI Prescriptions No.#1 and No.#2.


### 2.2. Proof of PMI for the Master Kinetic Equation

The next step is to prove that the functionalIM(ρ1o(N),b) defined above (see (28)) indeed exhibits a monotonic time-decreasing behavior which is consistent with the MKI Prescriptions No.#3 and No.#4, which are realized respectively by(i)
the time derivative inequality (18) and the conditions of existence of kinetic equilibrium (21);(ii)
the validity of the inequalityIM(ρ1(N)(t))≤1.In order to reach the proofs of these properties let us preliminarily determine the variation across a binary collision occurring between particles1 and 2 of the total directional kinetic energy M(v1,v2,b) (see (25)), namely, the phase-space scalar function ΔM(v1,v2,b)≡M(v1(+),v2(+),b)-M(v1v2,b). One obtains(40)ΔMv1,v2,b=b·n12n12·v12+v12+·b-b·n122n12·v12+2,the rhs being expressed in terms of the outgoing particle velocities (v1(+),v2(+)) only. Then, the following proposition holds.Theorem 2 (property of macroscopic irreversibility (master equation PMI theorem)).
Let us assume thatρ1(N)(x1,t)≡ρ1(N)(r1,v1,t) is an arbitrary stochastic particular solution of the Master kinetic equation (A.4) with initial condition ρ1o(N)(x1) such that the integral KM(ρ1o(N)(x1),b) exists and is nonvanishing. Then it follows that(i)
PropositionP21: one finds that for all t≥to(41)∂∂tKMρ1Nt,b=-N-1σ2∫U11dv1∫U12dv2∫Ωdr1Θ¯1∂Ωr¯1∫-dΣ21v12+·n12b·n122n12·v12+2∂ρ^1Nr1,v1+,t∂r1·∂∂r2ρ^1Nr2=r1+σn21,v2+tk2Nr1,r2=r1+σn21,t≤0.(ii)
PropositionP22: the inequality (42)∂∂tIMρ1Nt,b≤0holds globally (i.e., identically for allt≥to) so that necessarily IM(ρ1(N)(t),b) is globally defined too, being also prescribed so that(43)IMρ1Nt,b≤1.(iii)
PropositionP23: one finds that a given time t∈I with t≥to(44)∂∂tKMρ1Nt,b=0⇔ρ1Nx1,t≡ρ1MNv1.Proof.
Consider first the proof of PropositionP21 which requires evaluation of the partial time derivative ∂/∂tKM(ρ1(N)(t),b). Upon invoking the first form of the Master kinetic equation (see (A.1) in Appendix A), explicit differentiation of KM(ρ1(N)(t),b) delivers(45)∂∂tKMρ1Nt,b=-∫Γ11dx1Mv1,bΘ¯1∂Ωr¯1k1Nr1,t-v1·∂∂r1∂2ρ^1Nx1,t∂r1·∂r1-∫Γ11dx1Mv1,bΘ¯1∂Ω∂2ρ^1Nx1,t∂r1·∂r1∂∂tk1Nr1,t, namely, upon integration by parts in the first integral on the rhs,(46)∂∂tKMρ1Nt,b=-∫Γ11dx1Mv1,bΘ¯1∂Ω∂2ρ^1Nx1,t∂r1·∂r1∂∂t+v1·∂∂r1k1Nr1,t. Hence, thanks to the differential identity (B.12) it follows(47)∂∂tKMρ1Nt,b=-N-1∫Γ11dx1Mv1,bΘ¯1∂Ωr¯1∂2ρ^1Nx1,t∂r1·∂r1∫Γ¯12dx2v12·n12×δr1-r2-σk2Nr1,r2,tρ^1Nx2,t. Performing an integration by parts with respect to r1 and upon invoking the first differential identity (B.14) reported in Appendix B one obtains therefore(48)∂∂tKMρ1Nt,b=WMρ1Nt,b,(49)WMρ1Nt,b≡∫Γ11dx1Mv1,bΘ¯1∂Ωr¯1∂ρ^1Nx1,t∂r1N-1×∫Γ¯12dx2v12·n12∂∂r1δr1-r2-σk2Nr1,r2,tρ^1Nx2,t,where ∂/∂r1δ(r1-r2-σ)=-∂/∂r2δ(r1-r2-σ). Hence performing a further integration by parts with respect to r2 and using the second differential identity on (B.14) (see Appendix B) the previous equation finally yields(50)WMρ1Nt,b=N-1∫Γ11dx1Θ¯1∂Ωr¯1∫Γ¯12dx2v12·n12Mv1,v2,b×δr1-r2-σk2Nr1,r2,t∂ρ^1Nx1,t∂r1·∂ρ^1Nx2,t∂r2,where the symmetry property with respect to the exchange of states x1,x2 has been invoked. In the previous equation the integration on the Dirac delta can be performed at once letting(51)∫Γ11dx1Θ¯1∂Ω∫Γ12dx2δr1-r2-σ=σ2∫U11dv1∫U12dv2∫Ωdr1Θ¯1∂Ωr¯1∫+dΣ21v12·n12-∫-dΣ21v12·n12, where the solid-angle integrations in the two integrals on the rhs are performed, respectively, on the outgoing (+) and incoming (-) particles. Furthermore, it is obvious that thanks to the causal form of MCBC (see (C.3) in Appendix C) the integral on outgoing particles ∫(+)dΣ21 can be transformed to a corresponding integration on incoming ones, namely, ∫(-)dΣ21. Thus, the contributions in the two phase-space integrals only differ because of the variation ΔM(v1,v2,b) of the total directional kinetic energy of particles 1 and 2. This implies that(52)WMρ1Nt,b=N-1σ2∫U11dv1∫U12dv2∫Ωdr1Θ¯1∂Ωr¯1∫-dΣ21×v12·n12ΔMv1,v2,b∂ρ^1Nr1,v1+,t∂r1·∂ρ^1Nr2=r1+σn21,v2+,t∂r2×k2Nr1,r2=r1+σn21,t,where the solid-angle integration is performed on the incoming particles whereas ΔM(v1,v2,b) is evaluated in terms of the outgoing particles (+) and therefore must be identified with the second equation on the rhs of (40). Consider now the dependence in terms of the outgoing particle velocities v1(+) and v2(+) in the previous phase-space integral. The velocity dependence contained in the factors v12·n12 and ∂ρ^1(N)(r1,v1(+),t)/∂r1·∂ρ^1(N)(r2,v2(+)t)/∂r2 is symmetric with respect to the variables v1(+) and v2(+). On the other hand, as a whole, the same integral should remain unaffected with respect to the exchange of the outgoing particle velocities v1(+)⇔v2(+). This means that the only term in ΔM(v1,v2,b) which gives a (possibly) nonvanishing contribution is -b·n122n12·v12(+)2. As a consequence it is found that(53)∂∂tKMρ1Nt,b≡WMρ1Nt,b=-N-1σ2∫U11dv1∫U12dv2×∫Ωdr1Θ¯1∂Ωr¯1∫-dΣ21∂ρ^1Nr1,v1+,t∂r1·∂ρ^1Nr2=r1+σn21,v2+,t∂r2×v12+·n12b·n122n12·v12+2k2Nr1,r2=r1+σn21,t≤0,and hence ∂/∂tKM(ρ1(N)(t),b) is necessarily negative or null, the second case occurring only if ∂ρ^1(N)(r1,v1(+),t)/∂r1≡0 and consequently ∂ρ^1(N)(r2=r1+σn21,v2(+)t)/∂r2≡0 too.
The proof of PropositionP22 follows in a similar way. In fact, first, one notices that thanks to the global validity of the 1-body PDF [13] the 1-body PDF ρ1(N)(t) necessarily belongs to the functional class of stochastic PDFs ρ1(N)(x1,t) prescribed so thatalso the local characteristic scale length defined aboveLρ(t) (see (16))is larger than zero and finite. As a consequence it follows that both the functional KM(ρ1(N)(t),b) and IM(ρ1(N)(t),b) (see (22)) are globally defined too. Consider in fact the representation of KM(ρ1(N)(t),b) achieved in Theorem 1 and given by (35). Next, let us notice that thanks to (16) the characteristic scale length(54)Lμ,min≡inf⁡∂ρ^1Nx1,t∂r1-1 is necessarily strictly positive. Then, upon noting that k2(N)(r1,r2,t)≤1 and ∫Γ12dx2δr2-r1-σρ^1N(x2,t)≤sup⁡n^1(N)(r2,t)<+∞, with n^1(N)(r2,t) being the velocity moment n^1(N)(r2,t)=∫U1(2)dv2ρ^1(N)(r2,v2,t), it follows that(55)KMρ1Nt,b≤N-1Lμ,minsup⁡n^1Nr2,t∫Γ11dx1Θ¯1∂Ωr¯1Mv1,bρ^1Nx1,t, where the integral on the rhs is necessarily bounded. This happens because ρ1(N)(x1,t) belongs to the functional class ρ1(N)(t) and therefore n^1(N)(r2,t) is bounded, while, at the same time, the phase-space moments indicated above necessarily exist. Furthermore, since ∂/∂tIM(ρ1(N)(t),b)≡1/KMo∂/∂tKM(ρ1(N)(t),b), the inequality (53) implies (42) and (43) too. Finally, since ρ^1(N)(r1,v1,t) is a solution of the Master kinetic equation ∂/∂r1ρ^1(N)(x1,t)≡0 occurs if and only if ρ^1(N)(x1,t) coincides with a Maxwellian kinetic equilibrium of the type (6). This result proves therefore also Proposition P23.The implication of Theorem2 is therefore that provided the initial value KM(ρ1o(N)(x1),b) is nonvanishing then necessarily:(i)
the functionalKM(ρ1(N)(t),b) is monotonically decreasing and thus KM(ρ1(N)(t),b)≤KM(ρ1o(N)(x1),b);(ii)
similarly the MKI functionalIM(ρ1(N)(t),b) is monotonically decreasing too, i.e., IM(ρ1(N)(t),b)≤IM(ρ1o(N)(x1),b);(iii)
bothKM(ρ1(N)(t),b) and IM(ρ1(N)(t),b) are nonnegative.


### 2.3. Proof of the DKE Property for the Master Kinetic Equation

Let us now show that in validity of Theorems1 and 2 the time-evolved ρ1(N)(x1,t) necessarily must decay asymptotically for τ≡t-to→+∞ to kinetic equilibrium, i.e., that the limit function limτ→+∞⁡ρ1(N)(x1,t)≡ρ1∞(N)(x1) exists and it necessarily coincides with a Maxwellian kinetic equilibrium of the type (6). In this regard the following proposition holds.Theorem 3 (asymptotic behavior ofIM(ρ1(N)(t),b) (master equation-DKE theorem)).
Let us assume that the initial conditionρ1o(N)(x1)∈ρ1o(N)(x1) is such that the corresponding functional KM(ρ1o(N)(x1),b) is nonvanishing, i.e., in view of Theorem 1 necessarily >0. Then it follows that the corresponding time-evolved solution of the Master kinetic equation ρ1(N)(x1,t) in the limit τ≡t-to→+∞ necessarily must decay to kinetic equilibrium, i.e.,(56)limτ→+∞⁡ρ1Nx1,t=ρ1MNv1.Proof.
In order to reach the thesis it is sufficient to prove that necessarily(57)limτ→+∞⁡∂∂tIMρ1Nt,b=0.In fact, let us assume “ad absurdum” that ∂/∂tIM(ρ1(N)(t),b)≤-k2 with k2>0 being a real constant. Then Theorem 2 (Proposition P22) requires that(58)limτ→+∞⁡IMρ1Nt,b≤-limτ→+∞⁡t-tok2=-∞, a result which contradicts Theorem 1. This proves the validity of (57). Furthermore, by construction ∂/∂tIM≡1/KMo∂/∂tKM and furthermore ∂/∂tKM is identified with the functional WM(ρ1(N)(t),b)≤0 which is determined by (53). At this point one notices that, thanks to continuity of the functional WM(ρ1(N)(t),b), the identity(59)limτ→+∞⁡∂∂tIMρ1Nx1,t,b=WMρ1∞Nx1,b holds, where, thanks to global existence of the 1-body PDF (see [13]), the limit function(60)limτ→+∞⁡ρ1Nx1,t≡ρ1∞Nx1 necessarily exists. As a consequence (57) requires also the equation(61)WMρ1∞Nx1,b=0 to hold. Upon invoking Proposition P23 of Theorem 2 this implies that necessarily ρ1∞(N)(x1)=ρ1M(N)(v1) so the thesis (56) is proved. Incidentally, thanks to Theorem 1, this requires also that(62)limτ→+∞⁡IMρ1Nt,b=IMρ1∞Nx1,b=0.


### 2.4. Remarks

A few remarks are worth being pointed out regarding the results presented above.(1)
Remark #1. The choice of the MKI functional considered here (see (22)) is just one of the infinite particular admissible realizations which meet the complete set of MKI prescriptions indicated above. In particular, the choice of the velocity moment M(v1,b) considered here (see (24)) remains in principle arbitrary, since v1·b2 can be equivalently replaced, for example, by any factor of the form v1·b2n, with n≥1. Furthermore it is obvious that M(v1,b) can be replaced by any function of the form M(v1,b)+ΔM(v1,b), being ΔM(v1,b) prescribed in such a way that its contribution to ∂/∂tIM vanishes identically so that the validity of the inequality (42) in Theorem 2 is preserved. This implies in turn that the prescription of the MKI functional IM(ρ1(N)(t)) remains in principle nonunique.(2)
Remark #2. A possible issue is related to the requirement that the renormalized 1-body PDF, as the 1-body PDF itself, are strictly positive at all times and are nonvanishing. Here it is sufficient to state that an elementary consequence of the theory of the Master kinetic equation developed in [3] is that, provided the corresponding initial N-body PDF set at a prescribed initial time to is strictly positive in the whole N-body phase-space, both the corresponding renormalized 1-body PDF as the 1-body PDF remain necessarily strictly positive globally in time too and everywhere in the 1-body phase-space.(3)
Remark #3. It must be stressed that the signature of the time derivative ∂/∂tIM(ρ1(N)(t),b) actually depends crucially on the adoption of the causal form of MCBC (i.e., see (C.1) or (C.3) in Appendix C) rather than the anticausal one (given instead by (C.2)). The first choice is mandatory in view of the causality principle. Indeed, it is immediate to prove that ∂/∂tIM(ρ1(N)(t),b) changes signature if the anticausal MCBC (C.2) is invoked.(4)
Remark #4. Theorem 2 warrants that macroscopic irreversibility, namely, the inequality ∂/∂tIM(ρ1(N)(t),b)≤0, occurs specifically because of (a) the time-variation of the b-directional total kinetic energy which occurs at arbitrary binary-collision events; (b) the occurrence of a velocity-space anisotropy in the 1-body PDF, i.e., the fact that the same PDF may not coincide with a local Maxwellian PDF.(5)
Remark #5.The existence of the limit function limt→+∞⁡ρ1(N)(x1,t)=ρ1∞(N)(x1) follows uniquely as a consequence of the global existence theorem holding for the Master kinetic equation [13].(6)
Remark #6.Last but not least, the fact that the same limit function may coincide or not with the Maxwellian kinetic equilibrium (6) depends crucially on the functional setting prescribed for the same PDF ρ1(N)(x1,t). More precisely, DKE can only occur provided ρ1(N)(x1,t) is a suitably smooth stochastic PDF such that the MKI functional exists for the corresponding initial PDF at time to, i.e., ρ1(N)(x1,to)=ρ1o(N)(x1).Theorems1–3 represent the main results reached in the paper of what may be referred to as thePMI/DKE theoryfor finite hard-sphere systems and which have concerned the axiomatic formulation in such a context of the notion of macroscopic irreversibility and the related one of decay to kinetic equilibrium.


## 2.1. Proof of the Nonnegativity of the MKI Information Measure

The strategy adopted for the proof of the MKI Prescriptions No.#1 and No.#2 is to show initially the validity of the information-measure left inequality in (17), namely, that IM(ρ1(N)(t),b) cannot acquire negative values for arbitrary t≥to. The result is established by the following theorem.Theorem 1 (nonnegativity ofKM(ρ1o(N)(x1),b), KM(ρ1(N)(t),b), and IM(ρ1(N)(t),b)).
Let us assume thatρ1(N)(x1,t) is an arbitrary stochastic and suitably smoothly differentiable, particular solution of the Master kinetic equation (A.4) prescribed so that the integral (23) expressed in terms of the initial PDF, namely, KM(ρ1o(N)(x1),b), is nonvanishing. Then, it follows necessarily that(i)
PropositionP11:(29)KMρ1oNx1,b>0.(ii)
PropositionP12: the corresponding time-evolved functional KM(ρ1(N)(t),b) for all t∈I with t>to is such that(30)KMρ1Nt,b≥0.(iii)
PropositionP13: for all t∈I with t>to the functional IM(ρ1(N)(t),b) fulfills the inequality(31)IMρ1Nt,b≥0.(iv)
PropositionP14: the following necessary and sufficient condition holds at a given time t∈I with t≥to:(32)KMρ1Nt,b=0⇔ρ1Nx1,t≡ρ1MNv1.Proof.
One first notices thatKM(ρ1(N)(t),b) can be equivalently written in the form(33)KMρ1Nt,b≡-∫Γ11dx1Θ¯1∂Ωr¯1Mv1,bk1Nr1,t∂2ρ^1Nx1,t∂r1·∂r1, where in order that the same functional exists it is obvious that the renormalized 1-body PDF ρ^1(N)(x1,t) must be of class C(2). Integrating by parts and noting that the gradient term ∂Θ¯1(∂Ω)(r¯)/∂r1 gives a vanishing contribution to the phase-space integral, this yields equivalently(34)KMρ1Nt,b≡∫Γ11dx1Θ¯1∂Ωr¯1Mv1,b∂k1Nr1,t∂r1·∂ρ^1Nx1,t∂r1. Therefore, upon invoking (B.9) reported in Appendix B, direct substitution delivers(35)KMρ1Nt,b=N-1∫Γ11dx1Θ¯1∂Ωr¯1Mv1,b∂ρ^1Nx1,t∂r1·∫Γ12dx2n12×δr2-r1-σΘ¯2∂Ωr¯2ρ^1Nx2,tk2Nr1,r2,t.Next, invoking the identity n12δr2-r1-σ=-∂/∂r2Θ¯r2-r1-σ and noting again that ∂/∂r2Θ¯2(∂Ω)(r¯) gives vanishing contribution, one can perform a further integration by parts with respect to r2. This permits to cast the rhs of previous equation in the form(36)KMρ1Nt,b≡KM1ρ1Nx1,t,b+ΔKM1ρ1Nx1,t,b.Here the two terms on the rhs of (36) are defined as follows: (1) the first term KM(1)(ρ1(N)(t),b) is symmetric and nonnegative, so that it can be expressed so to carry the total directional kinetic energy M(v1,v2,b) of particles 1 and 2 (see (25)). Hence, it takes the form(37)KM1ρ1Nx1,t,b=N-1∫Γ11dx1∫Γ12dx2Θ¯1∂Ωr¯1Θ¯2∂Ωr¯2×∂ρ^1Nx1,t∂r1·∂ρ^1Nx2,t∂r2k2Nr1,r2,tMv1,v2,bΘ¯r2-r1-σ.(2) The second term ΔKM(1)(ρ1(N)(x1,t),b) reads instead(38)ΔKM1ρ1Nx1,t,b≡N-1∫Γ11dx1Θ¯1∂Ωr¯1∂ρ^1Nx1,t∂r1·∫Γ12dx2×Θ¯2∂Ωr¯2Mv1,bΘ¯r2-r1-σρ^1Nx2,t∂∂r2k2Nr1,r2,t,where ∂/∂r2k2(N)(r1,r2,t) is given by the differential identity (B.10) reported in Appendix B. Thus, upon invoking the identity n23δr3-r2-σ=-∂/∂r3Θ¯r3-r2-σ, one notices that an integration by parts can be performed also with respect to r3. This means that a procedure analogous to the one used for the calculation of ∂k1(N)(r1,t)/∂r1 can be invoked and iterated at all orders, i.e., up to the (N-1)-body occupation coefficient (see (B.11) in Appendix B). As a consequence the functional KM(ρ1(N)(t),b) can be represented in terms of a finite sum of the form KM(ρ1(N)(t),b)≡∑j=1,N-1KM(j)(ρ1(N)(x1,t),b) in which each term of the sum KM(j)(ρ1(N)(x1,t),b) is nonnegative and symmetric. This implies therefore that the same functional KM(ρ1(N)(t),b) can be cast in the form(39)KMρ1Nt,b=∫Γ11dx1∫Γ12dx2Θ¯1∂Ωr¯1Θ¯2∂Ωr¯2Mv1,v2,b×Fr1,r2,t∂ρ^1Nx1,t∂r1·∂ρ^1Nx2,t∂r2Θ¯r2-r1-σ,with M(v1,v2,b)≥0 being the total directional kinetic energy (25) and F(r1,r2,t) a suitable real scalar kernel which is symmetric in the variables r1 and r2. Hence KM(ρ1(N)(t),b) actually defines a nonnegative functional. This proves the validity of the inequality (31) (Proposition P11).
In a similar way also the remaining propositions can be established. In fact, invoking (28) it follows that the inequalities (30) and (31)—and hence also Propositions P12 and P13—manifestly hold too. Finally, regarding the proof of Proposition P14, one notices that KM(ρ1(N)(t),b)≡0 if and only if identically ∂/∂r1ρ^1(N)(x1,t)≡0. Since ρ^1(N)(x1,t) is by construction a solution of the Master kinetic equation it follows that this requires necessarily that ρ^1(N)(x1,t) must coincide with the local Maxwellian ρ1M(N)(v1) (see (6)) and hence (32) must hold too under the same realization (Proposition P14).The conclusion is therefore that the definition of the MKI functional (22) given above in terms of KM(ρ1(N)(t),b) and KMo (see (23)) is indeed consistent with the physical prerequisites represented by the MKI Prescriptions No.#1 and No.#2.


## 2.2. Proof of PMI for the Master Kinetic Equation

The next step is to prove that the functionalIM(ρ1o(N),b) defined above (see (28)) indeed exhibits a monotonic time-decreasing behavior which is consistent with the MKI Prescriptions No.#3 and No.#4, which are realized respectively by(i)
the time derivative inequality (18) and the conditions of existence of kinetic equilibrium (21);(ii)
the validity of the inequalityIM(ρ1(N)(t))≤1.In order to reach the proofs of these properties let us preliminarily determine the variation across a binary collision occurring between particles1 and 2 of the total directional kinetic energy M(v1,v2,b) (see (25)), namely, the phase-space scalar function ΔM(v1,v2,b)≡M(v1(+),v2(+),b)-M(v1v2,b). One obtains(40)ΔMv1,v2,b=b·n12n12·v12+v12+·b-b·n122n12·v12+2,the rhs being expressed in terms of the outgoing particle velocities (v1(+),v2(+)) only. Then, the following proposition holds.Theorem 2 (property of macroscopic irreversibility (master equation PMI theorem)).
Let us assume thatρ1(N)(x1,t)≡ρ1(N)(r1,v1,t) is an arbitrary stochastic particular solution of the Master kinetic equation (A.4) with initial condition ρ1o(N)(x1) such that the integral KM(ρ1o(N)(x1),b) exists and is nonvanishing. Then it follows that(i)
PropositionP21: one finds that for all t≥to(41)∂∂tKMρ1Nt,b=-N-1σ2∫U11dv1∫U12dv2∫Ωdr1Θ¯1∂Ωr¯1∫-dΣ21v12+·n12b·n122n12·v12+2∂ρ^1Nr1,v1+,t∂r1·∂∂r2ρ^1Nr2=r1+σn21,v2+tk2Nr1,r2=r1+σn21,t≤0.(ii)
PropositionP22: the inequality (42)∂∂tIMρ1Nt,b≤0holds globally (i.e., identically for allt≥to) so that necessarily IM(ρ1(N)(t),b) is globally defined too, being also prescribed so that(43)IMρ1Nt,b≤1.(iii)
PropositionP23: one finds that a given time t∈I with t≥to(44)∂∂tKMρ1Nt,b=0⇔ρ1Nx1,t≡ρ1MNv1.Proof.
Consider first the proof of PropositionP21 which requires evaluation of the partial time derivative ∂/∂tKM(ρ1(N)(t),b). Upon invoking the first form of the Master kinetic equation (see (A.1) in Appendix A), explicit differentiation of KM(ρ1(N)(t),b) delivers(45)∂∂tKMρ1Nt,b=-∫Γ11dx1Mv1,bΘ¯1∂Ωr¯1k1Nr1,t-v1·∂∂r1∂2ρ^1Nx1,t∂r1·∂r1-∫Γ11dx1Mv1,bΘ¯1∂Ω∂2ρ^1Nx1,t∂r1·∂r1∂∂tk1Nr1,t, namely, upon integration by parts in the first integral on the rhs,(46)∂∂tKMρ1Nt,b=-∫Γ11dx1Mv1,bΘ¯1∂Ω∂2ρ^1Nx1,t∂r1·∂r1∂∂t+v1·∂∂r1k1Nr1,t. Hence, thanks to the differential identity (B.12) it follows(47)∂∂tKMρ1Nt,b=-N-1∫Γ11dx1Mv1,bΘ¯1∂Ωr¯1∂2ρ^1Nx1,t∂r1·∂r1∫Γ¯12dx2v12·n12×δr1-r2-σk2Nr1,r2,tρ^1Nx2,t. Performing an integration by parts with respect to r1 and upon invoking the first differential identity (B.14) reported in Appendix B one obtains therefore(48)∂∂tKMρ1Nt,b=WMρ1Nt,b,(49)WMρ1Nt,b≡∫Γ11dx1Mv1,bΘ¯1∂Ωr¯1∂ρ^1Nx1,t∂r1N-1×∫Γ¯12dx2v12·n12∂∂r1δr1-r2-σk2Nr1,r2,tρ^1Nx2,t,where ∂/∂r1δ(r1-r2-σ)=-∂/∂r2δ(r1-r2-σ). Hence performing a further integration by parts with respect to r2 and using the second differential identity on (B.14) (see Appendix B) the previous equation finally yields(50)WMρ1Nt,b=N-1∫Γ11dx1Θ¯1∂Ωr¯1∫Γ¯12dx2v12·n12Mv1,v2,b×δr1-r2-σk2Nr1,r2,t∂ρ^1Nx1,t∂r1·∂ρ^1Nx2,t∂r2,where the symmetry property with respect to the exchange of states x1,x2 has been invoked. In the previous equation the integration on the Dirac delta can be performed at once letting(51)∫Γ11dx1Θ¯1∂Ω∫Γ12dx2δr1-r2-σ=σ2∫U11dv1∫U12dv2∫Ωdr1Θ¯1∂Ωr¯1∫+dΣ21v12·n12-∫-dΣ21v12·n12, where the solid-angle integrations in the two integrals on the rhs are performed, respectively, on the outgoing (+) and incoming (-) particles. Furthermore, it is obvious that thanks to the causal form of MCBC (see (C.3) in Appendix C) the integral on outgoing particles ∫(+)dΣ21 can be transformed to a corresponding integration on incoming ones, namely, ∫(-)dΣ21. Thus, the contributions in the two phase-space integrals only differ because of the variation ΔM(v1,v2,b) of the total directional kinetic energy of particles 1 and 2. This implies that(52)WMρ1Nt,b=N-1σ2∫U11dv1∫U12dv2∫Ωdr1Θ¯1∂Ωr¯1∫-dΣ21×v12·n12ΔMv1,v2,b∂ρ^1Nr1,v1+,t∂r1·∂ρ^1Nr2=r1+σn21,v2+,t∂r2×k2Nr1,r2=r1+σn21,t,where the solid-angle integration is performed on the incoming particles whereas ΔM(v1,v2,b) is evaluated in terms of the outgoing particles (+) and therefore must be identified with the second equation on the rhs of (40). Consider now the dependence in terms of the outgoing particle velocities v1(+) and v2(+) in the previous phase-space integral. The velocity dependence contained in the factors v12·n12 and ∂ρ^1(N)(r1,v1(+),t)/∂r1·∂ρ^1(N)(r2,v2(+)t)/∂r2 is symmetric with respect to the variables v1(+) and v2(+). On the other hand, as a whole, the same integral should remain unaffected with respect to the exchange of the outgoing particle velocities v1(+)⇔v2(+). This means that the only term in ΔM(v1,v2,b) which gives a (possibly) nonvanishing contribution is -b·n122n12·v12(+)2. As a consequence it is found that(53)∂∂tKMρ1Nt,b≡WMρ1Nt,b=-N-1σ2∫U11dv1∫U12dv2×∫Ωdr1Θ¯1∂Ωr¯1∫-dΣ21∂ρ^1Nr1,v1+,t∂r1·∂ρ^1Nr2=r1+σn21,v2+,t∂r2×v12+·n12b·n122n12·v12+2k2Nr1,r2=r1+σn21,t≤0,and hence ∂/∂tKM(ρ1(N)(t),b) is necessarily negative or null, the second case occurring only if ∂ρ^1(N)(r1,v1(+),t)/∂r1≡0 and consequently ∂ρ^1(N)(r2=r1+σn21,v2(+)t)/∂r2≡0 too.
The proof of PropositionP22 follows in a similar way. In fact, first, one notices that thanks to the global validity of the 1-body PDF [13] the 1-body PDF ρ1(N)(t) necessarily belongs to the functional class of stochastic PDFs ρ1(N)(x1,t) prescribed so thatalso the local characteristic scale length defined aboveLρ(t) (see (16))is larger than zero and finite. As a consequence it follows that both the functional KM(ρ1(N)(t),b) and IM(ρ1(N)(t),b) (see (22)) are globally defined too. Consider in fact the representation of KM(ρ1(N)(t),b) achieved in Theorem 1 and given by (35). Next, let us notice that thanks to (16) the characteristic scale length(54)Lμ,min≡inf⁡∂ρ^1Nx1,t∂r1-1 is necessarily strictly positive. Then, upon noting that k2(N)(r1,r2,t)≤1 and ∫Γ12dx2δr2-r1-σρ^1N(x2,t)≤sup⁡n^1(N)(r2,t)<+∞, with n^1(N)(r2,t) being the velocity moment n^1(N)(r2,t)=∫U1(2)dv2ρ^1(N)(r2,v2,t), it follows that(55)KMρ1Nt,b≤N-1Lμ,minsup⁡n^1Nr2,t∫Γ11dx1Θ¯1∂Ωr¯1Mv1,bρ^1Nx1,t, where the integral on the rhs is necessarily bounded. This happens because ρ1(N)(x1,t) belongs to the functional class ρ1(N)(t) and therefore n^1(N)(r2,t) is bounded, while, at the same time, the phase-space moments indicated above necessarily exist. Furthermore, since ∂/∂tIM(ρ1(N)(t),b)≡1/KMo∂/∂tKM(ρ1(N)(t),b), the inequality (53) implies (42) and (43) too. Finally, since ρ^1(N)(r1,v1,t) is a solution of the Master kinetic equation ∂/∂r1ρ^1(N)(x1,t)≡0 occurs if and only if ρ^1(N)(x1,t) coincides with a Maxwellian kinetic equilibrium of the type (6). This result proves therefore also Proposition P23.The implication of Theorem2 is therefore that provided the initial value KM(ρ1o(N)(x1),b) is nonvanishing then necessarily:(i)
the functionalKM(ρ1(N)(t),b) is monotonically decreasing and thus KM(ρ1(N)(t),b)≤KM(ρ1o(N)(x1),b);(ii)
similarly the MKI functionalIM(ρ1(N)(t),b) is monotonically decreasing too, i.e., IM(ρ1(N)(t),b)≤IM(ρ1o(N)(x1),b);(iii)
bothKM(ρ1(N)(t),b) and IM(ρ1(N)(t),b) are nonnegative.


## 2.3. Proof of the DKE Property for the Master Kinetic Equation

Let us now show that in validity of Theorems1 and 2 the time-evolved ρ1(N)(x1,t) necessarily must decay asymptotically for τ≡t-to→+∞ to kinetic equilibrium, i.e., that the limit function limτ→+∞⁡ρ1(N)(x1,t)≡ρ1∞(N)(x1) exists and it necessarily coincides with a Maxwellian kinetic equilibrium of the type (6). In this regard the following proposition holds.Theorem 3 (asymptotic behavior ofIM(ρ1(N)(t),b) (master equation-DKE theorem)).
Let us assume that the initial conditionρ1o(N)(x1)∈ρ1o(N)(x1) is such that the corresponding functional KM(ρ1o(N)(x1),b) is nonvanishing, i.e., in view of Theorem 1 necessarily >0. Then it follows that the corresponding time-evolved solution of the Master kinetic equation ρ1(N)(x1,t) in the limit τ≡t-to→+∞ necessarily must decay to kinetic equilibrium, i.e.,(56)limτ→+∞⁡ρ1Nx1,t=ρ1MNv1.Proof.
In order to reach the thesis it is sufficient to prove that necessarily(57)limτ→+∞⁡∂∂tIMρ1Nt,b=0.In fact, let us assume “ad absurdum” that ∂/∂tIM(ρ1(N)(t),b)≤-k2 with k2>0 being a real constant. Then Theorem 2 (Proposition P22) requires that(58)limτ→+∞⁡IMρ1Nt,b≤-limτ→+∞⁡t-tok2=-∞, a result which contradicts Theorem 1. This proves the validity of (57). Furthermore, by construction ∂/∂tIM≡1/KMo∂/∂tKM and furthermore ∂/∂tKM is identified with the functional WM(ρ1(N)(t),b)≤0 which is determined by (53). At this point one notices that, thanks to continuity of the functional WM(ρ1(N)(t),b), the identity(59)limτ→+∞⁡∂∂tIMρ1Nx1,t,b=WMρ1∞Nx1,b holds, where, thanks to global existence of the 1-body PDF (see [13]), the limit function(60)limτ→+∞⁡ρ1Nx1,t≡ρ1∞Nx1 necessarily exists. As a consequence (57) requires also the equation(61)WMρ1∞Nx1,b=0 to hold. Upon invoking Proposition P23 of Theorem 2 this implies that necessarily ρ1∞(N)(x1)=ρ1M(N)(v1) so the thesis (56) is proved. Incidentally, thanks to Theorem 1, this requires also that(62)limτ→+∞⁡IMρ1Nt,b=IMρ1∞Nx1,b=0.


## 2.4. Remarks

A few remarks are worth being pointed out regarding the results presented above.(1)
Remark #1. The choice of the MKI functional considered here (see (22)) is just one of the infinite particular admissible realizations which meet the complete set of MKI prescriptions indicated above. In particular, the choice of the velocity moment M(v1,b) considered here (see (24)) remains in principle arbitrary, since v1·b2 can be equivalently replaced, for example, by any factor of the form v1·b2n, with n≥1. Furthermore it is obvious that M(v1,b) can be replaced by any function of the form M(v1,b)+ΔM(v1,b), being ΔM(v1,b) prescribed in such a way that its contribution to ∂/∂tIM vanishes identically so that the validity of the inequality (42) in Theorem 2 is preserved. This implies in turn that the prescription of the MKI functional IM(ρ1(N)(t)) remains in principle nonunique.(2)
Remark #2. A possible issue is related to the requirement that the renormalized 1-body PDF, as the 1-body PDF itself, are strictly positive at all times and are nonvanishing. Here it is sufficient to state that an elementary consequence of the theory of the Master kinetic equation developed in [3] is that, provided the corresponding initial N-body PDF set at a prescribed initial time to is strictly positive in the whole N-body phase-space, both the corresponding renormalized 1-body PDF as the 1-body PDF remain necessarily strictly positive globally in time too and everywhere in the 1-body phase-space.(3)
Remark #3. It must be stressed that the signature of the time derivative ∂/∂tIM(ρ1(N)(t),b) actually depends crucially on the adoption of the causal form of MCBC (i.e., see (C.1) or (C.3) in Appendix C) rather than the anticausal one (given instead by (C.2)). The first choice is mandatory in view of the causality principle. Indeed, it is immediate to prove that ∂/∂tIM(ρ1(N)(t),b) changes signature if the anticausal MCBC (C.2) is invoked.(4)
Remark #4. Theorem 2 warrants that macroscopic irreversibility, namely, the inequality ∂/∂tIM(ρ1(N)(t),b)≤0, occurs specifically because of (a) the time-variation of the b-directional total kinetic energy which occurs at arbitrary binary-collision events; (b) the occurrence of a velocity-space anisotropy in the 1-body PDF, i.e., the fact that the same PDF may not coincide with a local Maxwellian PDF.(5)
Remark #5.The existence of the limit function limt→+∞⁡ρ1(N)(x1,t)=ρ1∞(N)(x1) follows uniquely as a consequence of the global existence theorem holding for the Master kinetic equation [13].(6)
Remark #6.Last but not least, the fact that the same limit function may coincide or not with the Maxwellian kinetic equilibrium (6) depends crucially on the functional setting prescribed for the same PDF ρ1(N)(x1,t). More precisely, DKE can only occur provided ρ1(N)(x1,t) is a suitably smooth stochastic PDF such that the MKI functional exists for the corresponding initial PDF at time to, i.e., ρ1(N)(x1,to)=ρ1o(N)(x1).Theorems1–3 represent the main results reached in the paper of what may be referred to as thePMI/DKE theoryfor finite hard-sphere systems and which have concerned the axiomatic formulation in such a context of the notion of macroscopic irreversibility and the related one of decay to kinetic equilibrium.


## 3. Consistency of MPI/DKE Theory with Microscopic Dynamics

The crucial problem which arises in the context of the ab initio theory is in some sense analogous to that occurring in the Boltzmann and Grad kinetic theories. The question is in fact whether these phenomena are actually consistent with the fundamental symmetry properties of the underlying Boltzmann-Sinai CDS. The problem posed in the present section concerns, more precisely, the consistency with the time-reversible, energy-conserving, evolution of the underlyingN-body Boltzmann-Sinai classical dynamical system SN-CDS.(1)
First issue: consistency with the microscopic reversibility principle. This is related to the famous objection raised by Loschmidt to the Boltzmann equation and Boltzmann H-theorem: i.e.,whether and possibly alsohow it may be possible to reconcile the validity of the reversibility principle for the SN-CDS with the manifestation of a decay of the 1-body PDF to kinetic equilibrium, i.e., the uniform Maxwellian PDF of the form (6), as predicted by the above Master equation-DKE Theorem. That a satisfactory answer to this question is actually possible follows from considerations which are based on the axiomatic (ab initio) statistical description realized by the Master kinetic equation. In this regard it is worth recalling the discussion reported above concerning the role of MCBC regarding the functional ∂/∂tIM(ρ1(N)(t)). In particular, it is obvious that the signature depends on whether the causal (or anticausal) form of MCBC is invoked (see Appendix C). Such a choice is not arbitrary since, for consistency with the causality principle, it must depend on the microscopic arrow of time, i.e., the orientation of the time axis chosen for the reference frame. Based on these premises, consistency between the occurrence of macroscopic irreversibility associated with the DKE phenomenon and the principle of microscopic reversibility can immediately be established. Indeed, it is sufficient to notice that when a time-reversal or a velocity-reversal is performed on the SN −CDS the form of the collision boundary conditions (i.e., in the present case the MCBC provided by (C.1) in Appendix C) must be changed, replacing them with the corresponding anticausal ones, i.e., (C.2). This implies that MKI functional decreases in both cases, i.e., after performing the time-reversal, so that no contradiction can possibly arise in this case between Theorem 3 and the microscopic reversibility principle.(2)
Second issue: consistency with Poincaré recurrence theorem.Similar considerations concern the consistency with the recurrence theorem due to Poincaré as well as the conservation of total (kinetic) energy for the SN −CDS. In fact, first, as it follows from [3], by construction the Master collision operator admits the customary Boltzmann collisional invariants, including total kinetic energy of colliding particles. Hence, total energy conservation is again warranted for SN −CDS. Second, regarding Poincaré recurrence theorem, it concerns the Lagrangian phase-space trajectories of the SN −CDS, i.e., the fact that almost all of these trajectories return arbitrarily close—in a suitable sense to be prescribed in terms of a distance defined on the N-body phase-space—to their initial condition after a suitably large “recurrence time”. Incidentally, its magnitude depends strongly both on the same initial condition and the notion of distance to be established on the same phase-space. Nevertheless, such a “recurrence effect” influences only the Lagrangian time evolution of the N-body PDF which occurs along the same Lagrangian N-body phase-space trajectories. Instead, the same recurrence effect has manifestly no influence on the time evolution of the Eulerian 1-body PDF which is advanced in time in terms of the Eulerian kinetic equation represented by the Master kinetic equation. Therefore the mutual consistency of DKE and Poincaré recurrence theorem remains obvious.Hence, in the framework of the axiomatic ab initio theory based on the Master kinetic equation the full consistency is warranted with the microscopic dynamics of the underlying Boltzmann-Sinai CDS.


## 4. Physical Implications

Let us now investigate the physical interpretation and main implications emerging from the PMI/DKE theory developed here. The first issue is related to the physical mechanism at the basis of the PMI/DKE phenomenology.It is well known that in the context of Boltzmann kinetic theory the property of macroscopic irreversibility and the occurrence of the DKE phenomenon are both ascribed to the Boltzmann H-theorem, both in its original formulation [7] and in its modified form introduced by Boltzmann himself while attempting to reply [28] to Loschmidt objection [27] (see also [30, 31] together with different views on the matter given in [24, 26]). As recalled above, this is expressed in terms of the production rate for the Boltzmann-Shannon entropy ∂/∂tS(ρ1(t)), with S(ρ1(t)) being interpreted as a measure of the ignorance associated with a solution of the Boltzmann equation. In fact the customary interpretation is that they arise specifically because of the validity of the entropic inequality (10), i.e., the monotonic increase of S(ρ1(t)), and the corresponding entropic equality (11) stating a necessary and sufficient condition for kinetic equilibrium. Such a theorem is actually intimately related to the equation itself. In fact both the theorem and the equation generally hold only for stochastic PDFs ρ1(t)=ρ1(x1,t) which are suitably smooth and not for distributions [1]. According to Boltzmann’s original interpretation, however, both the Boltzmann equation and Boltzmann H-theorem should only hold when the so-called Boltzmann-Grad limit is invoked, i.e., based on the limit operator LBG≡limN→+∞Nσ2~O(1)⁡​ (see [4]).In striking departure from such a picture, one has the following:(i)
The axiomatic ab initio theory based on the Master kinetic equation and the present PMI/DKE theory are applicable to an arbitrary finite Boltzmann-Sinai CDS. This means that they hold for hard-sphere systems having a finite number of particles and with finite diameter and mass, i.e., without the need of invoking validity of asymptotic conditions.(ii)
The main departure with respect to Boltzmann kinetic theory arises because, as earlier discovered (see in particular the related discussion reported in [4]), the Boltzmann-Shannon entropy associated with an arbitrary stochastic 1-body PDF ρ1(N)(t)=ρ1(N)(x1,t) solution of the Master kinetic equation is identically conserved. Thus both PMI and DKE are essentially unrelated to the Boltzmann-Shannon entropy.(iii)
In the case of the Master kinetic equation the physical mechanism responsible for the occurrence of both PMI and DKE is unrelated to the Boltzmann-Shannon entropy. In fact, as shown here, it arises because of the properties of the MKI functionalIM(ρ1(N)(t),b) when it is expressed in terms of an arbitrary stochastic PDF ρ1(N)(t)=ρ1(N)(x1,t) solution of the Master kinetic equation. The only requirement is that the initial PDF ρ1oN(x1) is prescribed so that the corresponding MKI functional IM(ρ1o(N)(x1),b) exists.(iv)
As shown here the MKI functional is a suitably weighted phase-space moment ofρ1(N)(x1,t) which can be interpreted as an information measure for the same PDF, namely, belongs to the interval 0,1, and exhibits a monotonic-decreasing time-dependence, i.e., the property of macroscopic irreversibility.(v)
In addition bothIM(ρ1(N)(t),b) and its time derivative ∂/∂tIM(ρ1(N)(t),b) vanish identically if and only if the 1-body PDF coincides with a Maxwellian kinetic equilibrium of the type (6). This warrants in turn also the occurrence of the DKE phenomenon for ρ1(N)(x1,t), i.e., that for t-to→+∞ the same PDF must decay to a Maxwellian kinetic equilibrium of this type.(vi)
Finally, it is interesting to point out the peculiar behavior of the MKI functionalIM(ρ1(N)(t),b) and its time derivative ∂/∂tIM(ρ1(N)(t),b) when the Boltzmann-Grad limit is considered. In particular the 1- and 2-body occupation coefficients k1(N)(r1,t) and k2(N)(r1,r2,t) which appear in the Master kinetic equation (see Appendix B, (B.1) and (B.2)) become respectively(63)LBGk1Nr1,t=1LBGk2Nr1,r2,t=1.As a consequence the limit functionalsLBGIM(ρ1(N)(t),b) and LBG∂/∂tIM(ρ1(N)(t),b), are necessarily identically vanishing. This means that the present theory applies properly when the exact Master kinetic equation is considered and not to its asymptotic approximation obtained in the Boltzmann-Grad limit, namely the Boltzmann kinetic equation (see [3, 13]).An interesting issue, in the context of the PMI/DKE theory for the Master kinetic equation, is the role of MCBC in giving rise to the phenomena of macroscopic irreversibility and decay to kinetic equilibrium. Let us analyze for this purpose the two cases represented by unary and binary hard-sphere elastic collisions.First, let us recall the customary treatment of collision boundary conditions for unary collision events (also referred to as the so-called mirror reflection CBC; see for example Cercignani [6, 11]). This refers to the occurrence at a collision time ti of a single unary elastic collision for particle 1 at the boundary ∂Ω. Let us denote by n1 the inward normal to the stationary rigid boundary ∂Ω at the point of contact with the same particle and respectively x1(-)(t1)=r1(t1),v1(-)(t1) and x1(+)(t1)=r1(t1),v1(+)(t1) the incoming and outgoing particle states while v1(+) is determined by the elastic collision law for unary collisions, namely,(64)v1+=v1--2n1n1·v1-.Then, the PDF-conserving CBC for the 1-body PDF requires that the following identity holds:(65)ρ1Nx1+t1,ti=ρNx1-ti,ti,with ρ1N(x1(+)(t1),ti)≡ρ1N(+)(x1(+)(t1),ti) and ρN(x1(-)(ti),ti)≡ρN(-)(x1(-)(ti),ti) denoting the outgoing and incoming 1-body PDF, respectively. This identifies the PDF-conserving CBC usually adopted in Boltzmann kinetic theory [7] (Grad [10]; see also [2, 3]). The obvious physical implication of (65) is that ρ1N(x1(+)(t1),ti) (and ρN(x1(-)(ti),ti)) should be necessarily an even function of the velocity component n1·v1(-). Indeed as shown in [2, 3] the PDF-conserving CBC (65) should be replaced with a suitable CBC identified with the MCBC condition (see also Appendix C). When realized in terms of its causal form (predicting the outgoing PDF in terms of the incoming one) the MCBC for unary collisions is just(66)ρ1N+x1+t1,ti=ρN-x1+ti,ti, with ρN(-)(x1(+)(ti),ti) denoting the incoming 1-body PDF evaluated in terms of the outgoing state x1(+)(ti). Assuming left-continuity (see related discussion in [2]), this can then be identified with ρN(-)(x1(+)(ti),ti)≡ρN(x1(+)(ti),ti), thus yielding(67)ρ1N+x1+t1,ti=ρNx1+ti,ti.Equation (67) provides the physical prescription for the collision boundary condition, which is referred to as MCBC, holding for the 1-body PDF at arbitrary unary collision events. It is immediate to realize that the function ρN(x1(+)(ti),ti) needs not generally be even with respect to the velocity component n1·v1(-). In addition (67), just as (65), also permits the existence of the customary collisional invariants which in the case of unary collisions are X=1,n1·v1(-),v1·1__-n1n1, v12. As a consequence, one can show that (67) warrants at the same time also the validity of the so-called no-slip boundary conditions for the fluid velocity field V(r1,t) carried by the 1-body PDF ρ1N(x1,t).The treatment of MCBC holding for the2-body PDF in case of binary-collision events is analogous and is recalled for convenience in (C.2) of Appendix C.Let us briefly analyze the qualitative physical implications of (67) and (C.2) as far as the DKE theory is concerned. First, we notice that unary collisions cannot produce in a proper sense a velocity-isotropization effect since, as shown by (67), in such a case MCBC gives rise only to a change in the velocity distribution occurring during a unary collision due to a single component of the particle velocity, namely, n1·v1(-). As a consequence, this explains why unary collisions do not affect the rate of change of the MKI functional (see Theorem 2). Second, (C.2) shows—on the contrary—that binary collisions actually do affect by means of MCBC a velocity-spreading for the 1- and 2-body PDF. In particular, since the spreading effect occurs in principle for all components of particle velocities affecting both particles 1 and 2, this explains why binary collisions are actually responsible for the irreversible time evolution of the MKI functional (see Theorems 2 and 3).In turn, as implied by Theorem3, DKE arises because of the phenomenon of macroscopic irreversibility (Theorem 2). The latter arises due specifically to the possible occurrence of a velocity-space anisotropy which characterizes the 1-body PDF when the same PDF differs locally from kinetic equilibrium. In turn, this requires also that the 1-body PDF belongs to the functional class of admissible stochastic PDFs ρ1(N)(x1,t). In difference to Boltzmann kinetic theory, however, the key physical role is actually ascribed to the MKI functional IM(ρ1(N)(t)) rather than the Boltzmann-Shannon entropy S1(ρ1(N)(t)). In fact, recalled above, the same functional remains constant in time once the Master kinetic equation is adopted. Rather, as shown by Theorem 2, it is actually the Master kinetic information IM(ρ1(N)(t)) which exhibits the characteristic signatures of macroscopic irreversibility.The key differences arising between the two theories, i.e., the Boltzmann equation-DKE and the Master equation-DKE, are of course related to the different and peculiar intrinsic properties of the Boltzmann and Master kinetic equations. In particular, as discussed at length elsewhere (see [1, 3]), precisely because the Boltzmann equation is only an asymptotic approximation of the Master kinetic equation explains why a loss of information occurs in Boltzmann kinetic theory and consequently the related Boltzmann-Shannon entropy is not conserved.The present investigation shows that in the context of the Master kinetic equation, the macroscopic irreversibility property, i.e., the monotonic time-decay behavior of the MKI functional, can be explained at a more fundamental level, i.e., based specifically on the time-variation of theb-directional total kinetic energy which occurs at arbitrary binary-collision events.The Master equation-DKE theorem (Theorem3) given above provides a first-principle proof of the existence of the phenomenon of DKE occurring for the kinetic description of a finite number of extended hard-spheres, i.e., described by means of the Master kinetic equation. More precisely, the DKE phenomenon affects the 1-body PDFs belonging to the admissible functional class ρ1(N)(x1,t) determined according to Theorem 1 and requiring also that the local characteristic scale length Lρ(t) associated with ρ1(N)(x1,t) is nonzero at all times.


## 5. Conclusions

In this paper the problem of the property of microscopic irreversibility (PMI) and decay to kinetic equilibrium (DKE) of the1-body PDF has been addressed. In doing so original ideas and methods are adopted of the new ab initio theory for hard-sphere systems recently developed in the context of classical statistical mechanics [1, 2].These are not just small deviations from standard literature approaches. Such developments, in fact, have opened up a host of exciting new subjects of investigation and theoretical challenges in kinetic theory which arise thanks to, or in the context of, the ab initio approach to kinetic theory. Both are based in particular on the discovery of the Master kinetic equation first reported in [3], equation which has been adopted also in the present paper.The ab initio theory and specifically the present paper represent the attempt to reach a new foundational basis and axiomatic physical description of the classical statistical mechanics for hard-sphere systems. The topic which has been pursued here—which represents also a challenging test of the ab initio theory itself—concerns the investigation of the physical origins of PMI and the related DKE phenomenon arisingin finiteN-body hard-sphere systems. These issues refer in particular to(i)
the proof of the nonnegativity of Master kinetic information (Theorem1, Section 2.1) together with the property of macroscopic irreversibility (PMI; Theorem 2, Section 2.2);(ii)
the establishment of Theorem3 (Section 2.3) and the related proof of the property of decay to kinetic equilibrium (DKE);(iii)
the consistency of PMI and DKE with microscopic dynamics (Section3);(iv)
the analysis of the main physical implications of DKE (Section4).The theory presented here departs in several respects from previous literature and notably from Boltzmann kinetic theory. The main differences actually arise because of the nonasymptotic character of the new theory, i.e., the fact that it applies to arbitrary dense or rarefied systems for which the finite number and size of the constituent particles is accounted for [3]. In this paper basic consequences of the new theory have been investigated which concern the phenomenon of decay to global kinetic equilibrium.The present results are believed to be crucial, besides in mathematical research, for the physical applications of the ab initio theory statistical theory, i.e., the Master kinetic equation. Indeed, regarding challenging future developments of the theory one should mention among others the following examples of possible (and mutually related) routes worth to be explored. One is related to the investigation of the possible effects due to arbitrarily prescribed, i.e., nonvanishing, initial (binary or multibody)phase-space statistical correlations. As recalled above, in fact, the Master equation is appropriate only when suitably prescribed configuration-space statistical correlations are taken into account. The second goal concerns the investigation of the time-asymptotic properties of the same kinetic equation, for which the present paper may represent a useful basis. The third goal refers to the possible extension of the theory to mixtures formed by hard-spheres of different masses and diameter which possibly undergo both elastic and anelastic collisions. Finally, the fourth one concerns the investigation of hydrodynamic regimes for which a key prerequisite is provided by the DKE theory established here.



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