On the Quasistatic Limit of Dynamic Evolutions for a Peeling Test in Dimension One
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Abstract
The aim of this paper is to study the quasistatic limit of a onedimensional model of dynamic debonding. We start from a dynamic problem that strongly couples the wave equation in a timedependent domain with Griffith’s criterion for the evolution of the domain. Passing to the limit as inertia tends to zero, we find that the limit evolution satisfies a stability condition; however, the activation rule in Griffith’s (quasistatic) criterion does not hold in general, thus the limit evolution is not rateindependent.
Keywords
Griffith’s criterion Dynamic debonding Quasistatic limit Rateindependent evolution Singular perturbation Vanishing inertiaMathematics Subject Classification
35L05 35B40 35Q74 35R35 74J40 74K351 Introduction
In models that predict the growth of cracks in structures, it is often assumed that the process is quasistatic. The quasistatic hypothesis is that inertial effects can be neglected since the time scale of the external loading is very slow, or equivalently the speed of the internal oscillations is very large if compared with the speed of loading. The resulting evolutions are rateindependent, i.e. the system is invariant under time reparametrisation.
Starting from the scheme proposed in Francfort and Marigo (1998), quasistatic crack growth has been extensively studied in the mathematical literature. The existence of quasistatic evolutions in fracture mechanics has been proved in several papers concerning globally minimising evolutions (Dal Maso and Toader 2002; Chambolle 2003; Francfort and Larsen 2003; Dal Maso et al. 2005; Dal Maso and Zanini 2007; Dal Maso and Lazzaroni 2010; Cagnetti and Toader 2011; Lazzaroni 2011; Crismale et al. 2016) and vanishingviscosity solutions (Negri and Ortner 2008; Cagnetti 2008; Knees et al. 2008, 2010; Lazzaroni and Toader 2011; Artina et al. 2017; Almi 2017; Crismale and Lazzaroni 2016). We refer to Bourdin et al. (2008) for a presentation of the variational approach to fracture and to Mielke and Roubíček (2015) for the relations with the abstract theory of rateindependent systems. These results also show that quasistatic evolutions may present phases of brutal crack growth (appearing as time discontinuities in the quasistatic scale). In order to study fast propagations of cracks, a dynamical analysis is needed, since inertial effects have to be accounted for.
On the other hand, in the case of dynamic fracture, only preliminary existence results were given (Nicaise and Sändig 2007; Dal Maso and Larsen 2011; Dal Maso and Lucardesi 2016; Dal Maso et al. 2016). The main difficulty is that the equations of elastodynamics for the displacement have to be satisfied in a timedependent domain (i.e. the body in its reference configuration, minus the growing crack), while the evolution of the domain is prescribed by a firstorder flow rule. The resulting PDE system is strongly coupled, as in other models of damage or delamination [see, e.g. Frémond and Nedjar (1996), Bonetti et al. (2005), Bonetti and Bonfanti (2008), Rocca and Rossi (2014, 2015), Heinemann and Kraus (2015b, 2015a) for viscous flow rules and Roubíček (2009, 2010, 2013a, b), Larsen et al. (2010), Rossi and Roubíček (2011), Bartels and Roubíček (2011), Babadjan and Mora (2015), Lazzaroni et al. (2014), Roubíček and Tomassetti (2015), Maggiani and Mora (2016) and Rossi and Thomas (2016) for rateindependent evolutions of internal variables].
In few cases, it has been shown that the quasistatic hypothesis is a good approximation, that is, the dynamic solutions converge to a rateindependent evolution as inertia tends to zero. This was proved in Roubíček (2013a) and Lazzaroni et al. (2014) for damage models, including a damping term in the wave equation, and in Dal Maso and Scala (2014) in the case of perfect plasticity. On the other hand, even in finite dimension there are examples of singularly perturbed secondorder potentialtype equations (where the inertial term vanishes and the formal limit is an equilibrium equation), such that the dynamic solutions do not converge to equilibria (Nardini 2017). In finite dimension, if the equations include a friction term whose coefficient tends to zero as inertia vanishes, then the dynamic evolutions converge to a solution of the equilibrium equation (Agostiniani 2012).
In this paper we develop a “vanishing inertia” analysis for a model of dynamic debonding in dimension one. More precisely, we consider a peeling test for a perfectly flexible thin film initially attached to a rigid substrate; the process is assumed to depend only on one of the two variables parametrising the film. This model was studied in Dumouchel et al. (2008), Lazzaroni et al. (2012) and Dal Maso et al. (2016) and, as already observed in Freund (1990, Section 7.4), it is related to dynamic fracture since it features a coupling between the wave equation, satisfied in the debonded part of the film, and a flow rule for the evolution of the debonding front.
The evolution of the debonding front \(\ell _\varepsilon \) is determined by a criterion involving the internal energy, i.e. the sum of the potential and the kinetic energy, cf. (2.8). Specifically, this criterion involves the dynamic energy release rate, which is defined as a (sort of) partial derivative of the internal energy with respect to the elongation of the debonded region. We refer to the following section for its definition and for details on its existence, which was proved in Dal Maso et al. (2016). In this introduction we only stress that the dynamic energy release rate at time t, denoted by \(G_\varepsilon (t)\), depends only on the debonding speed \(\dot{\ell }_\varepsilon (t)\) and on the values of \(u_\varepsilon (s,x)\) for \(s\le t\).
The existence of a unique solution \((u_\varepsilon , \ell _\varepsilon )\) in a weak sense was proved under suitable assumptions in Dal Maso et al. (2016), see also below. Notice the strong coupling between (1.1) and (1.2): indeed, the variable \(\ell _\varepsilon \) appears in the domain of (1.1a), while \(G_\varepsilon \) in (1.2) depends on \(u_\varepsilon \). This is typical of dynamic fracture, too.
In fact, the peeling test is closely related to fracture. The debonded part of the film, here parametrised on the interval \((0,\ell _\varepsilon (t))\), corresponds to the uncracked part of a body subject to fracture; both domains are monotone in time, though in opposite directions, increasing in our case, decreasing in the fracture problem. The debonding propagation \(t\mapsto \ell _\varepsilon (t)\) corresponds to the evolution of a crack tip. The debonding front \(\ell _\varepsilon (t)\) has the role of a free boundary just as a crack. However, notice that cracks are discontinuity sets for the displacement, where a homogeneous Neumann condition is satisfied since they are traction free; in contrast, in the peeling test the displacement is continuous at \(\ell _\varepsilon (t)\) because of the Dirichlet constraint (1.1c): the debonding front is a discontinuity set for the displacement derivatives and represents a free boundary between \(\{x: u_\varepsilon (x,s)=0 \text { for every } s\le t \}\) and \(\{x: u_\varepsilon (x,s)\ne 0 \text { for some } s\le t\}\).
In this work we perform an asymptotic analysis of (1.1) and (1.2) as \(\varepsilon \) tends to zero, i.e. we study the limit of the system for quasistatic loading. Some results in this direction were given in Dumouchel et al. (2008) and Lazzaroni et al. (2012) in the specific case of a piecewise constant toughness.
The existence of a unique solution \((u^\varepsilon ,\ell ^\varepsilon )\) to the coupled problem (1.3) and (1.4) for a fixed \(\varepsilon >0\) is guaranteed by Dal Maso et al. (2016, Theorem 3.5), provided the data are Lipschitz and the local toughness is piecewise Lipschitz; moreover it turns out that \(u^\varepsilon \) is Lipschitz in time and space and \(\ell ^\varepsilon \) is Lipschitz in time. (See also Theorem 2.3 below.) The strategy employed there to prove the existence result relies on the specific onedimensional setting of the model. Indeed, it is possible to write the solution \(u^\varepsilon \) of the wave equation (1.3a) in terms of a onedimensional function \(f^\varepsilon \); more precisely, \(u^\varepsilon (t,x)\) depends on \(f^\varepsilon (x\pm \varepsilon t)\) through the D’Alembert formula (2.4). On the other hand, the dynamic energy release rate \(G^\varepsilon \) can also be expressed as a function of \(f^\varepsilon \), so Griffith’s criterion (1.4) reduces to a Cauchy problem which has a unique solution.
In this work, in order to study the limit of the solutions \((u^\varepsilon ,\ell ^\varepsilon )\) as \(\varepsilon \rightarrow 0\), we use again the onedimensional structure of the model. First we derive an a priori bound for the internal energy, uniform with respect to \(\varepsilon \); to this end, it is convenient to write the internal energy in terms of \(f^\varepsilon \), see Proposition 3.1. The uniform bound allows us to find a limit pair \((u,\ell )\). More precisely, since the functions \(\ell ^\varepsilon \) are nondecreasing and \(\ell ^\varepsilon (T)<L\), Helly’s Theorem provides a subsequence \(\varepsilon _k\) such that \(\ell ^{\varepsilon _k}\) converges for every t to a (possibly discontinuous) nondecreasing function \(\ell \). On the other hand, the uniform bound on \(u^{\varepsilon _k}\) in \(L^2(0,T;H^1(0,L))\) guarantees the existence of a weak limit u. We call \((u,\ell )\) the quasistatic limit of \((u^\varepsilon ,\ell ^\varepsilon )\).
Condition (1.6a) is guaranteed by Helly’s Theorem. By passing to the limit in (1.4b), we also prove that (1.6b) holds. For this result we use again the D’Alembert formula for \(u^\varepsilon \) and find the limit f of the onedimensional functions \(f^\varepsilon \). In fact, \(\dot{f}\) turns out to be related to \(u_x\) through an explicit formula, as we see in Theorem 3.11, which is our second main result.
In contrast, (1.6c) is in general not satisfied. This was already observed in the earlier paper (Lazzaroni et al. 2012), which presents an example of dynamic solutions whose limit violates (1.6c). The singular behaviour of these solutions is due to the choice of a toughness with discontinuities. Indeed, when the debonding front meets a discontinuity in the toughness, a shock wave is generated. The interaction of such singularities causes strong highfrequency oscillations of the kinetic energy, which affects the limit as the wave speed tends to infinity.
In the present paper, we continue the discussion of this kind of behaviour by providing an explicit case where (1.6c) does not hold in the limit even if the local toughness is constant and the other data are smooth. (See Sect. 4 and Remark 4.2). In our new example, the initial conditions are not at equilibrium, in particular the initial position \(u_0\) is not affine in \((0,\ell _0)\). Therefore, due to the previous results, the quasistatic limit cannot satisfy the initial condition, i.e. it has a time discontinuity at \(t=0\). Moreover, our analysis of the limit evolution \((u,\ell )\) shows that the internal energy given through the initial conditions is not relaxed instantaneously; its effects persist in a time interval where the evolution does not satisfy (1.6c). The surplus of initial energy, instantaneously converted into kinetic energy, cannot be quantified in a purely quasistatic analysis. For this reason the usual quasistatic formulation (1.6) is not suited to describe the quasistatic limit of our dynamic process.
2 Existence and Uniqueness Results
In this section we provide an outline of the results of existence and uniqueness for the coupled problem (1.3) and (1.4) for fixed \(\varepsilon >0\), proved in Dal Maso et al. (2016). The only difference with respect to Dal Maso et al. (2016) is that the speed of sound is \(\frac{1}{\varepsilon }\) instead of 1.
Definition 2.1
We say that \(u^\varepsilon \in \widetilde{H}^1(\Omega ^\varepsilon )\) (resp. in \(u^\varepsilon \in H^1(\Omega ^\varepsilon _T)\)) is a solution to (2.1) if \(\varepsilon ^2u^\varepsilon _{tt}u^\varepsilon _{xx}=0\) holds in the sense of distributions in \(\Omega ^\varepsilon \) (resp. \(\Omega ^\varepsilon _T\)), the boundary conditions (2.1b) and (2.1c) are intended in the sense of traces and the initial conditions (2.1d) and (2.1e) are satisfied in the sense of \(L^2(0,\ell _0)\) and \(H^{1}(0,\ell _0)\), respectively.
Condition (2.1e) makes sense since \(u_x^\varepsilon \in L^2(0,T;L^2(0,\ell _0))\) and, by the wave equation, \(u_{xx}^\varepsilon , u_{tt}^\varepsilon \in L^2(0,T;H^{1}(0,\ell _0))\), therefore \(u_t^\varepsilon \in H^1(0,T;H^{1}(0,\ell _0)) \subset C^0([0,T];H^{1}(0,\ell _0))\). Arguing as in Dal Maso et al. (2016, Section 1), it is possible to uniquely solve (2.1) by means of the D’Alembert formula, as it is stated in the next proposition.
Proposition 2.2
The following existence and uniqueness result for the coupled problem (2.1) and (2.14) for fixed \(\varepsilon >0\) was proved in Dal Maso et al. (2016, Theorem 3.5, Remark 3.6). The case of a toughness depending also on the debonding speed is addressed in the subsequent paper (Lazzaroni and Nardini 2017), where we discuss existence, uniqueness, and quasistatic limit.
Theorem 2.3
Let \(T>0\), assume (2.2), and let the local toughness \(\kappa \) be as in (2.12). Then, there is a unique solution \((u^\varepsilon ,\ell ^\varepsilon ) \in C^{0,1}(\Omega ^\varepsilon _T)\times C^{0,1}([0,T])\) to the coupled problem (2.1) and (2.14). Moreover, there exists a constant \(L_T^\varepsilon \) satisfying \(\dot{\ell }^\varepsilon \le L_T^\varepsilon < \frac{1}{\varepsilon }\).
3 A Priori Estimate and Convergence
3.1 A Priori Bounds
Proposition 3.1
Proof
We need to estimate the last term in (3.2). To this end, we notice that it is sufficient to get a uniform bound for \(\dot{f}^\varepsilon \) in \(L^\infty \) as in (3.4). Then the conclusion readily follows from the bounds on the initial conditions and on the toughness.
Remark 3.2
3.2 Convergence of the Solutions
The a priori bound on the energy allows the passage to the limit in \(\ell ^\varepsilon \).
Proposition 3.3
Proof
The Hausdorff convergence of \(\mathrm {Graph}\,\ell ^{\varepsilon _k}\) to \(\mathrm {Graph}^*\ell \) will be used in the proof of Theorem 3.5. To prove that \(\mathrm {Graph}\,\ell ^{\varepsilon _k}\) converges to \(\mathrm {Graph}^*\ell \) in the sense of Hausdorff, in the following lemma we employ the equivalent notion of Kuratowski convergence, whose definition is recalled below.
Lemma 3.4
The sets \(\mathrm {Graph}\,\ell ^{\varepsilon _k}\) converge to \(\mathrm {Graph}^*\ell \) in the sense of Hausdorff.
Proof
 (i)
Let \((t,x) {\in } [0,T] \times [0,L]\) and let \((t_k,x_k) \in \mathrm {Graph}\,\ell ^{\varepsilon _k}\) be a sequence such that \((t_{k_n},x_{k_n}) \rightarrow (t,x)\) for some subsequence. Then, \((t,x) \in \mathrm {Graph}^*\ell \).
 (ii)
For every \((t,x) \in \mathrm {Graph}^*\ell \) there exists a whole sequence such that \((t_k,x_k) \in \mathrm {Graph}\,\ell ^{\varepsilon _k}\) and \((t_k,x_k) \rightarrow (t,x)\).
We now prove condition (ii). Let \((t,x) \in \mathrm {Graph}^*\ell \). Then, for every \(\eta >0\) we have \(\ell (t\eta ) \le x \le \ell (t+\eta )\). We claim that there is a sequence \(x_k \rightarrow x\) such that \(x_k \in [\ell ^{\varepsilon _k}(t{}\eta ),\ell ^{\varepsilon _k}(t{+}\eta )]\). Specifically, if \(\ell (t{}\eta )< x < \ell (t{+}\eta )\) we take \(x_k:=x\); if \(x= \ell (t{}\eta )\) we take \(x_k:= \ell ^{\varepsilon _k}(t{}\eta )\); if \(x = \ell (t{+}\eta )\) we take \(x_k := \ell ^{\varepsilon _k}(t{+}\eta )\); in each case by pointwise convergence we conclude that \(x_k\rightarrow x\). Then, by continuity and monotonicity of \(\ell ^{\varepsilon _k}\), there exists \(t_k \in [t{}\eta ,t{+}\eta ]\) such that \(\ell ^{\varepsilon _k}(t_k)=x_k\). We conclude by the arbitrariness of \(\eta \). \(\square \)
We now investigate on the limit behaviour of \(u^\varepsilon \). The next theorem shows that the limit displacement solves problem (1.5).
Theorem 3.5
Proof
3.3 Convergence of the Stability Condition
At this stage of the asymptotic analysis we have found a limit pair \((u,\ell )\) that describes the evolution of the debonding when the speed of the external loading tends to zero. We now investigate on the limit of Griffith’s criterion (2.11) and we question whether the limit pair \((u,\ell )\) satisfies the quasistatic version of this criterion, i.e. whether \((u,\ell )\) is a rateindependent evolution according to the definition below.
Definition 3.6
Remark 3.7
Notice that (3.19) and (3.20) do not prescribe the behaviour of the system at time discontinuities. In order to determine suitable solutions, additional requirements can be imposed, e.g. requiring that the total energy is conserved also after jumps in time.
Remark 3.8
We now consider the pair \((u,\ell )\) obtained in Proposition 3.3 and Theorem 3.5. We want to verify if \((u,\ell )\) satisfies Definition 3.6. First, we observe that by construction (cf. the application of Helly’s theorem in Proposition 3.3) \(t \mapsto \ell (t)\) is nondecreasing, thus (3.20a) automatically holds for a.e. t.
Next, we show that (3.20b) is satisfied. We first prove a few technical results.
Lemma 3.9
Let \(\Omega \) be a bounded domain in \(\mathbb {R}^N\) and \(g_n \rightarrow 1\) in measure, with \(g_n:\Omega \rightarrow \mathbb {R}\) equibounded. Then, \(g_n \rightarrow 1\) strongly in \(L^2(\Omega )\).
Proof
Lemma 3.10
Let \(\Omega \) be a bounded open interval, \(g_n :\Omega \rightarrow \mathbb {R}\) a sequence of functions such that \(g_n \rightarrow 1\) in measure and let \(\rho _n :\Omega \rightarrow \Omega \) such that \(\rho _n^{1}\) are equiLipschitz and \(\rho _n \rightarrow 1\) uniformly in \(\overline{\Omega }\). Then, \(g_n \circ \rho _n \rightarrow 1\) in measure.
Proof
Theorem 3.11
Assume (2.2), (2.12), and (3.1) and let \((u^{\varepsilon },\ell ^{\varepsilon })\) be the solution to the coupled problem (2.1) and (2.14). Let L and \(\varepsilon _k\) be as in Proposition 3.3. Then, for a.e. \(t \in [0,T]\) conditions (3.20a) and (3.20b) are satisfied.
Proof
Remark 3.12
To this end, besides (3.1) we assume that \(w^{\varepsilon _k}\) converges to w strongly in \(H^1(0,T)\), that \(u_0^{\varepsilon _k}\) converges to \(u_0\) strongly in \(H^1(0,\ell _0)\), and that \(\varepsilon u_1\) converges to 0 strongly in \(L^2(0,\ell _0)\), i.e. the initial kinetic energy tends to zero. Then, by (3.25) the lower semicontinuity of the potential energy ensures that \(\mathcal E_\mathrm{qs}(t;\ell ,w)\le \liminf _{k\rightarrow \infty }\mathcal E^{\varepsilon _k}(t;\ell ^{\varepsilon _k},w^{\varepsilon _k})\). Passing to the limit in (3.2) and using (3.22), we obtain an energy inequality; the opposite inequality derives from (3.20b) with arguments similar to Remark 3.7. We thus obtain (3.21) which is equivalent to the activation condition (at least in time intervals with no jumps).
However, conditions (3.25) and (3.21) may not hold in general, as shown in the example of the following section. The example shows that in general (2.11c) does not pass to the limit and (3.20c) is not satisfied, even in the case of a constant toughness.
4 Counterexample to the Convergence of the Activation Condition
We now show an explicit case where the convergence of (2.11c)–(3.20c) fails.
A first counterexample to the convergence of the activation condition was presented in Lazzaroni et al. (2012). In this case, the singular behaviour is due to the choice of a toughness with discontinuities. More precisely, in Lazzaroni et al. (2012) it is assumed that \(\kappa (x)=\kappa _1\) in \((\ell _1,\ell _1+\delta )\) and \(\kappa (x)=\kappa _2\) for \(x\notin (\ell _1,\ell _1+\delta )\), where \(\kappa _1<\kappa _2\), \(\ell _1>\ell _0\), and \(\delta \) is sufficiently small; this models a short defect of the glue between the film and the substrate.
4.1 Analysis of Dynamic Solutions
Notice that \(\dot{f}^\varepsilon \) is constant in every interval \((a^\varepsilon _{i1},a^\varepsilon _{i})\), for \(i=1,2,3\).
Remark 4.1
By (2.7), (4.2), and (4.11), the displacement’s derivatives are piecewise constant; in the (t, x) plane, their discontinuities lie on some shock waves originating from \((0,\ell _0/2)\) (where the initial datum has a kink), travelling backword and forward in the debonded film, and reflecting at boundaries; they are represented by thick dashed lines in Fig. 5. Notice that the lines originating from \((0,\ell _0)\), employed in the construction above and marked by thin dashed lines in Fig. 5, are not discontinuity lines, since \(\dot{f}^\varepsilon _{3i} = \dot{f}^\varepsilon _{3i+1}\) for every \(i\ge 1\). This is actually a consequence of the compatibility among \(u^\varepsilon _0\), \(u_1^\varepsilon \), and \(\dot{\ell }^\varepsilon \) at \((0,\ell _0)\), namely \(\dot{u}^\varepsilon _0(\ell _0)\dot{\ell }^\varepsilon (0)+u_1^\varepsilon (\ell _0)=0\). We refer to Dal Maso et al. (2016, Remark 1.12) for more details on the regularity of the solutions.
4.2 Limit for Vanishing Inertia
We now study the limit \(\ell \) of the evolutions \(\ell ^\varepsilon \) as \(\varepsilon \rightarrow 0\). Notice that the initial conditions are not at equilibrium; in particular the initial position \(u_0(x)\) is not of the form \(\big [\frac{w(0)}{\ell _0}x+w(0)\big ] \vee 0\). Because of (3.11), there must be a time discontinuity at \(t=0\), i.e. the limit displacement u jumps to an equilibrium configuration. Nonetheless, we will show that \(\ell \) is continuous even at \(t=0\). In order to determine \(\ell \), the main point is to study the limit evolution of the debonding during the first phase characterised by the “stop and go” process illustrated above. Afterwards, during the second phase, the evolution of the debonding will proceed at constant speed, given by \(\lim _{\varepsilon \rightarrow 0}\dot{\ell }^\varepsilon _1=1\).
Therefore, for \(t\ge e^2{+}1\) we have \(\ell (t)=t\). In the time interval \([e^2{+}1,+\infty )\), corresponding to the second phase, the quasistatic limit \(\ell \) is a rateindependent evolution in the sense of Definition 3.6, see also Remark 3.8.
Notice that during the first phase the quasistatic limit \(\ell \) does not satisfy (3.20c), thus it does not comply with the notion of rateindependent evolution given in Definition 3.6. Indeed, since the local toughness is constant, Remark 3.8 implies that a rateindependent evolution must be piecewise affine (with possible jumps); in contrast, (4.17) is not the equation of a line. This result is similar to the one obtained in Lazzaroni et al. (2012) with a discontinuous local toughness: here we showed that a singular behaviour can be observed even if the local toughness is constant.
Remark 4.2
We recall that the initial displacement \(u_0^\varepsilon \) chosen in (4.1a) has a kink at \(\frac{\ell _0}{2}=1\). In this section, we showed that the interaction between the two slopes generates the “stop and go” process, which gives as a result the convergence to an evolution that does not satisfy Definition 3.6. However, this singular behaviour can be obtained even for a smooth initial datum. Indeed, let us consider a regularisation of \(u_0^\varepsilon \), coinciding with the original profile outside of \((1\frac{\delta }{2},1+\frac{\delta }{2})\), where \(\delta \in (0,1)\) is fixed. As a consequence of this choice, the function \(\dot{\ell }^\varepsilon \) differs from (4.12) only in a portion of the order \(\varepsilon \delta \) of each interval \((b_i^\varepsilon ,b_{i+1}^\varepsilon )\). The resulting evolution of the debonding front \(\ell ^\varepsilon \) is smooth. However, in the limit we observe the same qualitative behaviour described above, due to the interaction of the different slopes of the initial datum. This shows that the singular behaviour is not due to the choice of a initial datum with a kink.
4.3 Analysis of the Kinetic Energy

the nonquasistatic phase, where Griffith’s quasistatic criterion fails in the limit, is characterised by the presence of a relevant kinetic energy (of order one as \(\varepsilon \rightarrow 0\), at each fixed time);

during such first phase, kinetic energy oscillates and is exchanged with potential energy at a timescale of order \(\varepsilon \);

overall, the total mechanical energy decreases and is transferred into energy dissipated in the debonding growth;

as time increases, the maximal oscillations of the kinetic energy decrease and approach zero as \(t\rightarrow e^2+1\), i.e. all of the kinetic energy is converted into potential and dissipated energy;

in the second (stable) phase, for \(t\ge e^2+1\), the kinetic energy is of order \(\varepsilon ^2\) and does not influence the limit behaviour of the debonding evolution as \(\varepsilon \rightarrow 0\).
5 Conclusion
In this paper we have studied a dynamic peeling test and its limit for slow loading (or, equivalently, for vanishing inertia). We have proved that the limit displacement is at equilibrium (i.e. it is affine, see Theorem 3.5), while the limit debonding evolution is nondecreasing and satisfies a stability condition, precisely the energy release rate is controlled by the local toughness (Theorem 3.11). In contrast, the activation condition (3.20c) in Griffith’s criterion is in general not satisfied in the limit as shown by the example in Sect. 4. In fact, in that case the quasistatic energy balance does not hold, because of the presence of a relevant amount of kinetic energy during a first phase [whose asymptotic limit is given by (4.17) and (4.21)].
Such phase features a strictly positive debonding acceleration. In fact, in fracture mechanics, it was observed that the equation of motion involves crack acceleration in finite size specimens, as that considered here (Marder 1991; Goldman et al. 2010). The emergence of crack inertia is due to waves generated by the crack that interact with the crack after bouncing on the boundary. A similar phenomenon is observed in our example for the peeling test. The role of debonding acceleration in the quasistatic limit will be matter of future investigation: this may lead to a better characterisation of the notion of solution found in the limit.
Our example highlights the relevance of dynamical effects in debonding propagation under quasistatic loading: the quasistatic approximation given by Griffith’s criterion is not appropriate in this case since the kinetic energy can not be neglected. A similar behaviour was observed in Lazzaroni et al. (2012), arising from toughness defects. Our example shows a new situation where convergence to a rateindependent solution fails, due to initial data out of equilibrium. The same phenomenon is observed if we choose as initial condition \(u^\varepsilon _0(x):=u^\varepsilon ((a^\varepsilon _{3i+1}+a^\varepsilon _{3i+2})/2,x)\) (with the notation of the previous section), thus we obtain a nonquasistatic propagation even starting from an initial datum arbitrarily close to equilibrium.
Hence, our new example indicates that, in order to get convergence to Griffith’s criterion, one should essentially consider the trivial case of an initial datum at equilibrium. Indeed, also in finitedimensional singularly perturbed secondorder potentialtype equations, convergence to equilibrium is enforced by taking initial conditions at equilibrium (Nardini 2017); however, choosing an initial datum at equilibrium is not needed if such equations include a viscosity term tending to zero as inertia vanishes (Agostiniani 2012). This suggests that, in the case of the peeling test, Griffith’s criterion may hold in the quasistatic limit if the dynamic equations are damped. We leave this question open for further research.
Notes
Acknowledgements
Open access funding provided by University of Vienna. The authors wish to thank Gianni Dal Maso for several interesting discussions. This work is part of the INdAMGNAMPA Project 2016 “Multiscale analysis of complex systems with variational methods.” When most of this research was carried out, GL was affiliated to SISSA and supported by the ERC Advanced Grant 290888 “Quasistatic and Dynamic Evolution Problems in Plasticity and Fracture”; he is now funded by the Austrian Science Fund (FWF) project P 27052.
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