2. Mathematical formulation

Consider an $(x,y)$-plane where a fluid with viscosity $\mu _1$ lies in the lower half-plane $y < 0$ (region 1) and another of viscosity $\mu _2$ lies in the upper half-plane $y> 0$ (region 2), as indicated schematically in figure 1. An infinite interface along $y=0$ having surface tension $\sigma (x,t)$ is assumed to be laden with surfactant with concentration $\varGamma (x,t)$. Following Crowdy (Reference Crowdy2020,  Reference Crowdy2021a), it is assumed that a capillary number based on a clean-flow surface tension value is negligible so that the normal stress balance can be ignored and the interface can be taken to remain undeflected from $y=0$.

Both fluids are taken to be in the Stokes regime so that the velocity fields ${\boldsymbol{u}}_j = (u_j,v_j)$ for $j=1,2$ satisfy

(2.1a,b)\begin{equation} \mu_j\,\nabla^2 {\boldsymbol{u}}_j = \boldsymbol{\nabla} p_j, \quad \boldsymbol{\nabla} \boldsymbol{\cdot} {\boldsymbol{u}}_j = 0, \end{equation}

in their respective regions, where $p_j(x,y,t)$ is the fluid pressure in region $j$. The boundary conditions on $y=0$ are

(2.2a–c)\begin{equation} u_1 = u_2, \quad v_1 = v_2= 0, \quad \mu_1\, \frac{\partial u_1 }{\partial y} - \mu_2\, \frac{\partial u_2 }{\partial y} ={-} \frac{\partial \sigma }{\partial x},\end{equation}

which ensure, respectively, that fluid velocities are continuous across the interface, that the latter is a streamline, and that any net fluid shear stress across the interface is balanced by a Marangoni stress caused by the surface tension variation. Since we are interested in stagnant caps, in the far field in both fluid regions, a pure extensional flow is assumed with strain rate $\dot \gamma _0$ of the form

(2.3)\begin{equation} (u_j,v_j) \sim (-\dot \gamma_0 x, \dot \gamma_0 y), \quad j=1,2.\end{equation}

This far-field behaviour is consistent with the velocity boundary conditions in (2.2a–c). At large distances from the origin, however, the large velocities associated with such a flow are not consistent with the assumption of Stokes flow, and corrections will be needed there; but the focus here is on the vicinity of the origin towards which, in the presence of such a linear extensional flow, any surfactant on the interface is expected to be swept. Then, in the absence of surface diffusion or surfactant exchange with the bulk, a stagnant cap of surfactant, i.e. an immobilized region of the interface, will form (Palaparthi et al. Reference Palaparthi, Papageorgiou and Maldarelli2006; Crowdy Reference Crowdy2021a). Indeed, Crowdy (Reference Crowdy2021a) has recently given an analytical description of the time-dependent formation of such stagnant caps in the single-phase scenario where the upper region is taken to be a constant pressure zone. Several previous theoretical studies have examined such a two-dimensional single-fluid system with a view to understanding the transition region at the edge of a stagnant cap (Harper Reference Harper1992) and surfactant spreading (Jensen Reference Jensen1995).

Suppose that $L$ is a typical length scale associated with the support of the surfactant on the interface set by initial conditions, and let $\varGamma _\infty$ be the maximum surface packing concentration. According to (1.3), the Langmuir equation of state is

(2.4)\begin{equation} \sigma= \sigma_c + \beta \varGamma_\infty \log (1-\varGamma/\varGamma_\infty), \quad \beta = RT.\end{equation}

As already mentioned, a capillary number based on $\sigma _c$ is sufficiently small that surface deflection can be ignored and the interface assumed to be flat, occupying the line $y=0$ at all times (Crowdy Reference Crowdy2020,  Reference Crowdy2021a).

The surfactant is taken to be soluble to the upper fluid phase but insoluble to the lower fluid phase. The bulk concentration of surfactant molecules in region 2 is denoted by $c(x,y,t)$ and is governed by the advection diffusion equation

(2.5)\begin{equation} \frac{\partial c }{\partial t} +{\boldsymbol{u}}_2\boldsymbol{\cdot} \boldsymbol{\nabla} c = D\,\nabla^2 c(x,y,t), \end{equation}

where $D$ is the diffusion coefficient of the surfactant molecules in region 2. Since the aim is to study the possible dynamical effects of small amounts of ambient surfactant molecules not fed by any far-field source, it is assumed that any non-zero concentration of surfactant is localized in the sense that

(2.6)\begin{equation} c(x,y,t) \to 0 \quad {\rm as}\ |\boldsymbol{x}| \to \infty. \end{equation}

To represent surfactant going on and off the interface, we introduce the kinetic flux $j$ of surfactant onto the interface as

(2.7)\begin{equation} j = b c_s (\varGamma_\infty -\varGamma) - a \varGamma,\end{equation}

where $a$ is the desorption coefficient, $b$ is the adsorption coefficient, and

(2.8)\begin{equation} c_s = c(x,0,t) \end{equation}

is the so-called sublayer concentration (Palaparthi et al. Reference Palaparthi, Papageorgiou and Maldarelli2006). Conservation of surfactant requires that

(2.9)\begin{equation} D \left.\frac{\partial c }{\partial y} \right|_{y=0} = j.\end{equation}

On the interface at $y=0$, the surfactant evolution equation has the form

(2.10)\begin{equation} \frac{\partial \varGamma }{\partial t} + \frac{\partial (\varGamma \mathcal{U}) }{\partial x} = D_s\,\frac{\partial^2 \varGamma }{\partial x^2} + j, \end{equation}

where $\mathcal {U}(x,t) = u(x,0,t)$ is the surface slip velocity, and $D_s$ is the surface diffusion coefficient. For consistency with (2.6), we also assume that the far-field interface is clean,

(2.11)\begin{equation} \varGamma(x,t) \to 0 \quad\text{as } |x| \to \infty,\end{equation}

which means that any surfactant on the interface is also localized.

3. Non-dimensionalization

The velocity scale given by

(3.1)\begin{equation} \mathcal{U}_0 = \frac{\beta \varGamma_\infty }{2 (\mu_1 + \mu_2)}\end{equation}

turns out to be the most convenient for this analysis; it is the natural two-phase analogue of the single-phase velocity scale used by Crowdy (Reference Crowdy2021a,Reference Crowdyb). It is also a natural choice of velocity scale when a given initial distribution of surfactant simply spreads on the interface in the absence of any external flow. Velocities will be scaled with respect to this quantity, lengths with respect to $L$, time with respect to $L/\mathcal {U}_0$, fluid pressures with respect to $\mathcal {U}_0 \mu _i/L$, bulk concentrations with respect to $c_0$, and surface concentrations with respect to $\varGamma _\infty$.

While $\mathcal {U}_0$ does not depend on the externally applied flow, an alternative velocity scale for this problem is $\mathcal {U}_0^*=\dot {\gamma }_0 L$. The latter could have been used in the non-dimensionalization, but as discussed further in § 11, $\mathcal {U}_0$ can be used to evaluate the magnitude of the external flow and hence key quantities such as the Biot and Péclet numbers.

The flux balance equation (2.9) gives

(3.2)\begin{equation} D\,\frac{\partial c }{\partial y} = b c_s (\varGamma_\infty -\varGamma) - a \varGamma,\end{equation}

which, using primes to denote non-dimensional quantities, is

(3.3)\begin{equation} \frac{D c_0 }{L}\,\frac{\partial c' }{\partial y'} =a \varGamma_\infty \left( \frac{b }{a}\,c_0 c_s' (1-\varGamma') - \varGamma' \right).\end{equation}

Equation (3.3) can be written as

(3.4)\begin{equation} \frac{D c_0 }{\mathcal{U}_0 \varGamma_\infty}\, \frac{\partial c' }{\partial y'} =\frac{a L }{\mathcal{U}_0} \left( \frac{b }{a}\,c_0 c_s' (1-\varGamma') - \varGamma' \right), \quad {\rm on}\ y=0. \end{equation}

After dropping primes on all variables, (3.4) becomes

(3.5)\begin{equation} \alpha \left.\frac{\partial c }{\partial y} \right|_{y=0} = Bi ( k c_s (1-\varGamma) - \varGamma ),\end{equation}

where we have introduced the non-dimensional parameters

(3.6a–d)\begin{equation} Bi = \frac{a L }{\mathcal{U}_0}, \quad k = \frac{b c_0 }{a}, \quad \alpha = \frac{D c_0 }{\mathcal{U}_0 \varGamma_\infty} = \frac{c_0 L }{\varGamma_\infty\,Pe}, \quad Pe = \frac{\mathcal{U}_0 L }{D}. \end{equation}

Note that the parameter $\alpha$ can also be written as $\alpha =({bc_0}/{a})({aL}/{b\varGamma _\infty })({1}/{Pe})=k\chi /Pe$, which is the form used by Palaparthi et al. (Reference Palaparthi, Papageorgiou and Maldarelli2006) and Wang et al. (Reference Wang, Papageorgiou and Maldarelli2002); see § 1 also. The parameter $Bi$ is the Biot number, and $Pe$ is the bulk Péclet number associated with region 2. The surfactant evolution equation (2.10) has the non-dimensional form

(3.7)\begin{equation} \left.\frac{\partial \varGamma }{\partial t} + \frac{\partial (\varGamma \mathcal{U}) }{\partial x} = \frac{1 }{Pe_s}\,\frac{\partial^2 \varGamma }{\partial x^2} +\alpha\,\frac{\partial c }{\partial y} \right|_{y=0}, \end{equation}

where the surface Péclet number is

(3.8)\begin{equation} Pe_s = \frac{\mathcal{U}_0 L }{D_s}.\end{equation}

The far-field condition (2.3) becomes, in non-dimensional variables,

(3.9)\begin{equation} (u,v) \sim (-\dot \gamma x, \dot \gamma y),\end{equation}

where

(3.10)\begin{equation} \dot \gamma = \frac{L \dot \gamma_0 }{\mathcal{U}_0}. \end{equation}

Finally, the non-dimensional form of the advection–diffusion equation (2.5) is

(3.11)\begin{equation} Pe\,({\partial c \over \partial t}+ {\boldsymbol{u}}_2 \boldsymbol{\cdot} \boldsymbol{\nabla} c) = \nabla^2 c(x,y,t).\end{equation}

In the analysis to follow, a key assumption is that $Pe\ll 1$ so that $c(x,y,t)$ becomes a harmonic function in the bulk. This enables significant analytical gains and, most remarkably, the development of explicit solutions that can be evaluated readily. To justify this assumption, estimates of $Pe$ for physical systems describable by our analysis are provided in § 11. It is natural to expect the bulk and surface Péclet numbers to be of similar sizes, and in accordance with this, in § 7, analytical solutions (in terms of parabolic cylinder functions) are found for the steady states at arbitrary values of $Pe_s$, including for $Pe_s \ll 1$, which is most consistent with the bulk Péclet number assumption $Pe\ll 1$. An interesting feature of the mathematical analysis here is that it does not require small $Pe_s$, and it is therefore natural to explore theoretically what happens at much larger surface Péclet numbers, including the limiting case where $Pe_s \to \infty$, which happens to admit an even wider range of analytical possibilities. In § 8, the steady-state equilibria for $Pe_s \to \infty$ are studied in detail. It is shown there that parabolic cylinder functions are not needed to describe the equilibria in this case; instead, the solutions are found to involve only square root branch point singularities akin to those studied by Crowdy (Reference Crowdy2021a,Reference Crowdyb) for insoluble surfactants. Moreover, § 10 demonstrates that even the unsteady time evolution to these $Pe_s \to \infty$ equilibria can be captured in closed form using a complex generalization of the method of characteristics.

4. Fast reaction $Bi \to \infty$ in the dilute limit $\varGamma \ll 1$

Several approximations are now made that facilitate analytical progress and, at the same time, continue to describe a physically important regime. The first is that the levels of surfactant are low; this is the so-called dilute or gaseous limit, where $\varGamma \ll 1$. The second assumption is that the reaction kinetics are fast, corresponding to the limit $Bi \to \infty$. From the right-hand side of (3.5), it is clear that as $Bi \to \infty$, it must be true that

(4.1)\begin{equation} k c_s (1-\varGamma) - \varGamma = 0, \end{equation}

or, on rearrangement,

(4.2)\begin{equation} c_s = c(x,0,t) = \frac{\varGamma }{k (1 -\varGamma)} \approx \frac{\varGamma }{k},\end{equation}

where the dilute limit assumption $\varGamma \ll 1$ has been used. In this dilute limit, it is reasonable to use a linearized version of the equation of state (2.4), which in dimensional form is

(4.3)\begin{equation} \sigma(x,t) \approx \sigma_c - \beta\,\varGamma(x,t). \end{equation}

5. Complex variable formulation

It is shown in Crowdy (Reference Crowdy2021a,Reference Crowdyb) that for a linear equation of state, and at zero capillary number based on the clean-flow surface tension $\sigma _c$, the Stokes flow – as it happens, in either of the half-plane regions 1 or 2 – can be described by the streamfunctions

(5.1)\begin{equation} \psi_j(x,y,t) = {\rm Im} [(\bar{z} -z) f_j(z,t) ], \end{equation}

where, for now, it is helpful to revert to dimensional variables and note that the analytic functions $f_j(z,t)$ have the dimensions of a velocity. It is clear that both streamfunctions are constant on $y=0$, where $\bar {z}=z$, which ensures that the interface is a streamline. By following the formulation given in Crowdy (Reference Crowdy2021a,Reference Crowdyb), it can also be shown that

(5.2)\begin{equation} u_j = {\rm Re} [{-}2\,f_j(z,t)], \quad {\rm on}\ y=0, \quad j=1,2, \end{equation}

and that the dimensional Marangoni stress condition (2.2a–c) on $y=0$ can be written as

(5.3)\begin{equation} 2\mu_1 ({-}2\, {\rm Im} [f_1(z,t)]) - 2\mu_2 ({-}2 \,{\rm Im} [f_2(z,t)]) = \beta \varGamma, \end{equation}

where an integration with respect to $x$ has been carried out – see Crowdy (Reference Crowdy2021b). The velocity condition $u_1=u_2$ in (2.2a–c) then implies

(5.4)\begin{equation} {\rm Re} [f_1(z,t)] = {\rm Re} [f_2(z,t)] = {\rm Re} [\overline{f_2(z,t)}] = {\rm Re} [\overline{f_2}(z,t)],\end{equation}

where the Schwarz conjugate function $\bar {f}(z,t)$ of an analytic function $f(z,t)$ is defined by

(5.5)\begin{equation} \bar{f}(z,t) = \overline{f(\bar{z},t)}.\end{equation}

To satisfy the far-field strain condition (2.3), it is necessary that

(5.6)\begin{equation} f_j(z,t) \to \frac{\dot \gamma_0 z }{2} + {{O}}(1/z) \quad\text{as } |z| \to \infty, \quad j=1,2.\end{equation}

An important observation is that by the Schwarz reflection principle, if $f_2(z,t)$ is upper analytic with the far-field behaviour (5.6), then $\overline {f_2}(z,t)$ is a lower analytic function with the same far-field behaviour. Condition (5.4) therefore implies

(5.7)\begin{equation} {\rm Re} [f_1(z,t) - \overline{f_2}(z,t)] = 0, \quad y=0. \end{equation}

The function $f_1(z,t) - \overline {f_2}(z,t)$ is lower analytic and bounded in the far field, and has zero real part on $y=0$. It can be inferred immediately, by analytic continuation, that

(5.8)\begin{equation} f_1(z,t) = \overline{f_2}(z,t), \end{equation}

where condition (5.6) has been used to preclude the addition of a purely imaginary constant. The Marangoni condition (5.3) on $y=0$ can now be written as

(5.9)\begin{align} & 2\mu_1 ({-}2\, {\rm Im} [f_1(z,t)]) - 2\mu_2 ({-}2\, {\rm Im} [f_2(z,t)]) \nonumber \\ & \quad = 2\mu_1 ({-}2\, {\rm Im} [f_1(z,t)]) + 2\mu_2 ({-}2\,{\rm Im} [\overline{f_2}(z,t)])\nonumber\\ & \quad = 2(\mu_1+ \mu_2) ({-}2\, {\rm Im} [f_1(z,t)]) = \beta \varGamma, \end{align}

where the first equality follows from the elementary fact that the imaginary parts of complex conjugate quantities are negatives of each other. It is from (5.9) that the choice of velocity scale $\mathcal {U}_0$ in (3.1) emerges naturally. Therefore, using the scalings discussed previously and defining the non-dimensional analytic function $h(z,t)$ via

(5.10)\begin{equation} f_1(z,t) ={-}\frac{\mathcal{U}_0}{2}\,h(z,t), \end{equation}

it follows from (5.2) and (5.9) that

(5.11)\begin{equation} h(z,t) = \mathcal{U} + {\rm i} \varGamma \quad {\rm on}\ y=0, \end{equation}

where, for convenience, the quantity $\mathcal {U}(x,t)$ has been introduced to denote the common (non-dimensional) slip velocity of both fluids on the interface. Note that as $|z| \to \infty$ in region 1,

(5.12)\begin{equation} h(z,t) \to - \dot \gamma z + {{O}}(1/z). \end{equation}

At the same time, since the bulk Péclet number in region 2 is small, to leading order in this parameter, the diffusive concentration field is harmonic in region 2, and the upper analytic function

(5.13)\begin{equation} q(z,t) = c(x,y,t) + {\rm i}\,\varPi(x,y,t), \quad y > 0, \end{equation}

can be introduced, where $\varPi (x,y,t)$ is the harmonic conjugate to $c(x,y,t)$, and without loss of generality it is assumed that $\varPi (x,y,t) \to 0$ as $|z|\to \infty$. For consistency with (2.6), the condition

(5.14)\begin{equation} q(z,t) = {{O}}(1/z) \quad\text{as } |z| \to \infty\end{equation}

must hold.

6. A forced complex Burgers equation

This section follows the ideas developed in Crowdy (Reference Crowdy2021a,Reference Crowdyb). By the Cauchy–Riemann equations,

(6.1a,b)\begin{equation} \frac{\partial c }{\partial x} = \frac{\partial \varPi }{\partial y}, \quad \frac{\partial c }{\partial y} ={-}\frac{\partial \varPi }{\partial x}, \end{equation}

the non-dimensional surfactant evolution equation (3.7) can be written as

(6.2)\begin{equation} \frac{\partial \varGamma }{\partial t} + \frac{\partial (\varGamma \mathcal{U}) }{\partial x} + \alpha\,\frac{\partial \varPi }{\partial x} = \frac{1 }{Pe_s}\,\frac{\partial^2 \varGamma }{\partial x^2} \quad {\rm on}\ y=0.\end{equation}

In an important step, and following Crowdy (Reference Crowdy2021a,Reference Crowdyb), it is observed that this can be written as

(6.3)\begin{equation} {\rm Im} \left[ \frac{\partial h }{\partial t} + \frac{1 }{2}\,\frac{\partial h^2 }{\partial z} + \alpha\,\frac{\partial q }{\partial z} - \frac{1 }{Pe_s}\,\frac{\partial^2 h }{\partial z^2} \right] = 0 \quad {\rm on}\ y=0. \end{equation}

Another statement of the same condition is

(6.4)\begin{equation} {\rm Im} \left[ \frac{\partial h }{\partial t} + \frac{1 }{2}\,\frac{\partial h^2 }{\partial z} - \alpha\,\overline{\left(\frac{\partial q }{\partial z}\right)} - \frac{1 }{Pe_s}\, \frac{\partial^2 h }{\partial z^2} \right] = 0 \quad {\rm on}\ y=0, \end{equation}

where the fact that the imaginary parts of complex conjugate quantities are negatives of each other has been used. Since $\bar {z}=z$ on $y=0$, (6.4) is

(6.5)\begin{equation} {\rm Im} \left[ \frac{\partial h(z,t) }{\partial t} + h(z,t)\,\frac{\partial h(z,t) }{\partial z} - \alpha\,\frac{\partial \bar{q}(z,t) }{\partial z} - \frac{1 }{Pe_s}\,\frac{\partial^2 h(z,t) }{\partial z^2} \right] = 0 \quad {\rm on}\ y=0.\end{equation}

The quantity in square brackets is a lower analytic function, which, on use of (5.6) and (5.14), can be seen to tend to $\dot \gamma ^2 z+{{O}}(1/z)$ as $|z| \to \infty$. Then it follows, by analytic continuation off $y=0$, that

(6.6)\begin{equation} \frac{\partial h(z,t) }{\partial t} + h(z,t)\, \frac{\partial h(z,t) }{\partial z} - \alpha\, \frac{\partial \bar{q}(z,t) }{\partial z} - \frac{1 }{Pe_s}\,\frac{\partial^2 h(z,t)}{\partial z^2} = \dot \gamma^2 z. \end{equation}

This is a PDE that relates the two lower analytic functions $h(z,t)$ and $\bar {q}(z,t)$ everywhere in the complex plane (not just on the interface).

There is another relation between the two lower analytic functions $\bar {q}(z,t)$ and $h(z,t)$, however. The boundary condition (4.2) can be written as

(6.7)\begin{equation} {\rm Re} [q(z,t)] = {\rm Im} \left[\frac{1 }{k}\,h(z,t) \right] \quad {\rm on}\ y=0, \end{equation}

or equivalently as

(6.8)\begin{equation} {\rm Re} [q(z,t)] = {\rm Im} \left[\frac{1 }{k}\,(h(z,t)+ \dot \gamma z) \right]\quad {\rm on}\ y=0, \end{equation}

because the added term is zero. This, in turn, can be written as

(6.9)\begin{equation} {\rm Re} [q(z,t)] = {\rm Re} [\overline{q(z,t)}]= {\rm Re} [\bar{q}(z,t)] = {\rm Im} \left[ \frac{1 }{k}\,(h(z,t)+ \dot \gamma z ) \right]\quad {\rm on}\ y=0, \end{equation}

where the fact that a complex quantity and its complex conjugate have the same real part has been used, as well as the fact that $\bar {z}=z$ on $y=0$. Hence

(6.10)\begin{equation} {\rm Re} [\bar{q}(z,t)] ={-}{\rm Re} \left[\frac{{\rm i} }{k}\, (h(z,t)+ \dot \gamma z ) \right]\quad {\rm on}\ y=0. \end{equation}

But both $\bar {q}(z,t)$ and $-{\rm i} (h(z,t)+\dot \gamma z)/k$ are lower analytic functions that decay like $1/z$. Therefore, again by analytic continuation, we infer that

(6.11)\begin{equation} \bar{q}(z,t) ={-}\frac{{\rm i} }{k}\,(h(z,t) + \dot \gamma z),\end{equation}

where a possible additive purely imaginary constant has been omitted because both sides of (6.11) decay like $1/z$ as $|z| \to \infty$. Equation (6.11), which also holds everywhere, can now be used in (6.6) to give a PDE purely for $h(z,t)$:

(6.12)\begin{equation} \frac{\partial h }{\partial t} + h\,\frac{\partial h }{\partial z} + {\rm i} \delta \left( \frac{\partial h }{\partial z} + \dot \gamma \right) = \frac{1 }{Pe_s}\,\frac{\partial^2 h }{\partial z^2} + \dot \gamma^2 z, \end{equation}

where, for convenience, the important quantity

(6.13)\begin{equation} \delta = \frac{\alpha }{k},\end{equation}

has been introduced. Once (6.12) is solved for $h(z,t)$, the function $\bar {q}(z,t)$ – and hence its Schwarz conjugate $q(z,t)$, which describes the surfactant concentration in region 2 – follows from (6.11).

In the next section, the PDE (6.12) will be studied directly. Its nonlinearity is characteristic of the complex Burgers equation. Actually, by the change of independent variables from $(z,t)$ to $(Z,T)$ given by

(6.14)\begin{equation} Z = z- {\rm i}\delta t, \quad T=t, \end{equation}

and letting $H(Z,T)=h(z,t)$, (6.12) transforms to

(6.15)\begin{equation} \frac{\partial H }{\partial T} + H\,\frac{\partial H }{\partial Z} = \frac{1 }{Pe_s}\, \frac{\partial^2 H }{\partial Z^2} + \dot \gamma^2 (Z+{\rm i}\delta T) - {\rm i}\delta T,\end{equation}

which is precisely a forced complex Burgers equation for $H(Z,T)$ akin to that derived, in a different but closely related context, by Crowdy (Reference Crowdy2021a,Reference Crowdyb). Crowdy (Reference Crowdy2021b) showed, by introducing a change of dependent variables that is a complex generalization of the classical Cole–Hopf transformation, that PDEs of this kind can be linearized. It follows immediately from those arguments that the present problem can also be linearized in the same way. To sketch out the details, introduce a change of dependent variable given by

(6.16)\begin{equation} H(Z,T) ={-} \frac{2 }{Pe_s}\,\frac{\partial \log \varPhi(Z,T) }{\partial Z}; \end{equation}

then, after some algebra, it can be shown that $\varPhi$ satisfies

(6.17)\begin{equation} \frac{\partial \varPhi }{\partial T} = \frac{1 }{Pe_s}\,\frac{\partial^2 \varPhi }{\partial Z^2} - \frac{Pe_s }{2} \left(\dot \gamma^2 \left(\frac{Z^2 }{2} +{\rm i}\delta TZ \right) - {\rm i}\delta TZ \right) \varPhi. \end{equation}

The important point is that this is a linear PDE for the complex-valued function $\varPhi (Z,T)$. When $\dot \gamma = \delta =0$, it reduces to the classical complex heat equation (Crowdy Reference Crowdy2021b). Therefore, the nonlinear problem under consideration here can, by this combination of changes of both independent and dependent variables, be linearized at arbitrary surface Péclet number.

It may be that analytical solutions to (6.17) can be found, akin to those discovered by Crowdy (Reference Crowdy2021b) when $\dot \gamma = \delta =0$. Presumably, numerical schemes aimed at solving the linearized PDE (6.17) can be devised that are likely to have advantages over those that tackle the nonlinear system in its original form. These matters are under investigation.

Finally, it is worth remarking on some important differences between the complex PDE of Burgers type in (6.12) and the more familiar Burgers equation for a real-valued field familiar to fluid dynamicists from compressible gas dynamics, for example. The latter equation has been studied extensively in the literature, but very few of the analytical results there carry over to (6.12). One reason is that $h(z,t)$ is not constrained to be real-valued on $y=0$; indeed, its imaginary part is precisely the surfactant concentration of primary interest here. A second reason is that $h(z,t)$ is required to be lower analytic (at least, initially) with, moreover, non-negative imaginary part on $y=0$. These additional stipulations mean that most results derived for the Burgers equation governing real-valued functions on $y=0$ are simply not relevant to (6.12). These differences are discussed in more detail by Crowdy (Reference Crowdy2021b).

7. Steady equilibria for arbitrary $Pe_s$

In this section, steady equilibria of (6.12) are determined for arbitrary finite values of $Pe_s$. The nonlinear second-order ordinary differential equation (ODE) for $h(z)$ is

(7.1)\begin{equation} \frac{{\rm d} }{{\rm d} z}\left( \frac{h^2}{2}\right) + \mathrm{i}\delta \left( \frac{{\rm d} h}{{\rm d} z} +\dot{\gamma} \right) = \varepsilon\,\frac{{\rm d}^2 h}{{\rm d} z^2} + \dot{\gamma}^2 z, \end{equation}

where the parameter $\varepsilon =1/Pe_s$ has been introduced. Integration of (7.1) with respect to $z$ gives

(7.2)\begin{equation} h^2(z) + 2\mathrm{i}\delta (h(z) + \dot{\gamma} z) = 2 \varepsilon\,\frac{{\rm d} h}{{\rm d} z} + \dot{\gamma}^2 (z^2 - l^2), \quad l \in \mathbb{R}^+, \end{equation}

where a constant of integration has been chosen to ensure that as $\delta, \varepsilon \to 0$, the stagnant cap solution derived by Crowdy (Reference Crowdy2021a) is recovered, namely,

(7.3)\begin{equation} h_{SC}(z) ={-}\dot{\gamma}(z^2 - l^2)^{1/2}. \end{equation}

This solution corresponds to a stagnant cap of insoluble surfactant occupying the interval $[-l,l]$ on the real axis in the infinite surface Péclet number limit, or $\varepsilon \to 0$. Noticing that (7.2) is a first-order ODE of Riccati type, it can be converted into a linear second-order ODE by introducing

(7.4)\begin{equation} h(z) ={-}\frac{2\varepsilon}{w(z)}\,\frac{{\rm d} w}{{\rm d} z}. \end{equation}

(Harper (Reference Harper1992) made similar observations when studying the more limited steady-state transition from a semi-infinite shear-free interface over $x < 0$ to a semi-infinite immobilized interface over $x >0$.) This produces

(7.5)\begin{equation} \frac{{\rm d}^2 w}{{\rm d} z^2} - \frac{\mathrm{i}\delta}{\varepsilon}\, \frac{{\rm d} w}{{\rm d} z} - \frac{\dot{\gamma}^2 }{4\varepsilon^2} \left( z^2 -\frac{2\mathrm{i} \delta }{\dot{\gamma}}\,z - l^2 \right)w = 0. \end{equation}

A convenient transformation of independent variable from $z$ to $\zeta$ allows (7.5) to be rewritten as the following ODE for $\tilde {W}(\zeta ) \equiv w(z)$:

(7.6)\begin{equation} \frac{{\rm d}^2 \tilde{W}}{{\rm d} \zeta^2} - \frac{\mathrm{i} \delta }{\varepsilon}\,\frac{{\rm d} \tilde{W}}{{\rm d} \zeta} - \frac{\dot{\gamma}^2}{4\varepsilon^2} \left(\zeta^2 - l^2 + \frac{\delta^2}{\dot{\gamma}^2}\right) \tilde{W} = 0, \quad \zeta = z - \frac{\mathrm{i}\delta }{\dot{\gamma}}. \end{equation}

Geometrically, the $\zeta$-plane is the physical $z$-plane shifted down by a distance $\delta /\dot {\gamma }$. A second independent variable transformation from $\zeta$ to $\eta$ is also useful:

(7.7)\begin{equation} \eta = \mathrm{i} \left( \frac{\dot{\gamma}}{\varepsilon}\right)^{1/2} \zeta = \kappa z + \varDelta, \quad \kappa \equiv \mathrm{i} \left( \frac{\dot{\gamma}}{\varepsilon}\right)^{1/2}, \quad \varDelta \equiv \frac{\delta}{\sqrt{\dot{\gamma} \varepsilon}}. \end{equation}

Geometrically, the $\eta$-plane rescales the $\zeta$-plane and rotates it through angle ${\rm \pi} /2$. It transforms (7.6) into the following ODE for $W(\eta ) \equiv \tilde {W}(\zeta )$:

(7.8)\begin{equation} \frac{{\rm d}^2 W}{{\rm d} \eta^2} - \varDelta\,\frac{{\rm d} W}{{\rm d} \eta} - \left(\frac{\eta^2}{4} + \hat{a}\right) W(\eta) = 0, \quad \hat{a} \equiv \frac{l^2 \dot{\gamma}^2 - \delta^2}{4 \dot{\gamma} \varepsilon}. \end{equation}

This equation can be solved for $W(\eta )$ using the transformations (A1) and (A2) given in Appendix A, yielding the solution

(7.9)\begin{equation} W(\eta) = \exp\left( \frac{\varDelta \eta}{2} \right) U(a,\eta), \quad a \equiv \hat{a} + \frac{\varDelta^2}{4} = \frac{l^2\dot{\gamma}}{4\varepsilon} \geq 0, \end{equation}

where $U$ is the principal parabolic cylinder function. For more information on this function and its properties, the reader is referred to Appendix A and to Whittaker & Watson (Reference Whittaker and Watson1996), Olver et al. (Reference Olver, Olde Daalhuis, Lozier, Schneider, Boisvert, Clark, Mille, Saunders, Cohl and McClain2020) and Zhang & Jin (Reference Zhang and Jin1996). Note that the parameter $a$ in (7.9) should not be confused with the desorption kinetic constant introduced in (1.1). Solution (7.9) of the parabolic cylinder equation is the one satisfying the far-field condition (5.12), since

(7.10)\begin{equation} U(a, \eta) \sim \exp \left( -\frac{\eta^2}{4}\right) \eta^{{-}a-1/2} \quad \text{as } \lvert\eta \rvert \to \infty. \end{equation}

Substitution of (7.9) into (7.4) leads to

(7.11)\begin{equation} h(z) ={-}\frac{2\varepsilon \kappa}{W(\eta)}\,\frac{{\rm d} W}{{\rm d} \eta} ={-}2\varepsilon \kappa \left( \frac{\eta}{2} + \frac{\varDelta}{2} - \frac{U(a - 1, \eta)}{U(a, \eta)} \right), \end{equation}

where the identity

(7.12)\begin{equation} \frac{{\rm d} U(a, \eta)}{{\rm d} \eta} = \frac{1}{2}\,\eta\, U(a, \eta) - U(a-1,\eta) \end{equation}

has been used. Finally, use of (7.7) to rewrite all expressions in terms of the original independent variable $z$ leads to

(7.13)\begin{equation} h(z)= \dot{\gamma} z - 2\varepsilon \kappa \varDelta + 2 \varepsilon\kappa\,\frac{U(a -1,\kappa z+ \varDelta)}{U(a,\kappa z + \varDelta)}, \end{equation}

which, on use of (7.10), can be shown to exhibit the required far-field behaviour (5.12).

Some other checks on the solution (7.9) are required, however. The parabolic cylinder function $U(a,\eta )$ is an entire function of $\eta$, and hence of $z$, with an essential singularity at infinity (Olver et al. Reference Olver, Olde Daalhuis, Lozier, Schneider, Boisvert, Clark, Mille, Saunders, Cohl and McClain2020). In view of the logarithmic derivative of $w(z)$ in (7.4), in order to ensure the appropriate analyticity properties of (7.13), the zeros of $U(a,\eta )$ must be analysed. For non-negative values of $a$ – the only ones relevant here – $U(a,\eta )$ has no purely real or imaginary zeros, but infinitely many complex zeros, tending to infinity, that approach the conjugate rays $\text {Arg} (\eta ) = \pm 3{\rm \pi} /4$ as $\lvert \eta \rvert \to \infty$ (Olver et al. Reference Olver, Olde Daalhuis, Lozier, Schneider, Boisvert, Clark, Mille, Saunders, Cohl and McClain2020), where $\text {Arg}$ denotes the principal value of the argument function. These zeros lie entirely in the half-plane ${\rm Re}[\eta ] < 0$, and consequently, by unravelling the geometrical transformations associated with the aforementioned changes of independent variable as shown schematically in Appendix A, they lie in the upper half of the physical $z$-plane, shifted upwards by a distance $\delta /\dot {\gamma }$, and tending towards the rays $\text {Arg}(z-{\mathrm i}\delta /\dot {\gamma }) = {\rm \pi}/4, 3{\rm \pi} /4$ as $\lvert z\rvert \to \infty$. Indeed, the scaling given in (7.7) was chosen specifically to ensure this. The solution $h(z)$ is therefore lower analytic except for the required singular behaviour at infinity.

From (5.11), by evaluating the real and imaginary parts of (7.13) on $y =0$, the slip velocity $\mathcal {U}(x)$ and the surfactant concentration $\varGamma (x)$ can be extracted. Typical solutions are shown in figure 2 for varying $\varepsilon$ and $\delta$. As expected, as the effect of diffusion decreases (or $\varepsilon \to 0$) for low solubility (or $\delta \to 0$), the surfactant concentration tends towards the stagnant cap regime. This case is discussed in more detail in § 8. For a fixed surface Péclet number, or fixed $\varepsilon$, it is found that increasing the solubility parameter $\delta$ allows the surfactant to move on and off the interface at an increased exchange rate, mollifying the stagnant cap and hence remobilizing the interface. The combination of both increased solubility and diffusivity spreads the surfactant out along the interface while also allowing it to react faster with the bulk, resulting in reduced Marangoni stresses and a flow that behaves ultimately as if it is largely unaffected by the presence of surfactant.

Figure 2.

The effect of varying $\delta$ on (a,c,e) the slip velocity $\mathcal {U}(x)$, and (b,d, f) the surfactant concentration $\varGamma (x)$, with $\dot {\gamma } = l = 1$ as $\varepsilon \to 0$. Steady remobilization of the interface occurs as both the solubility and the diffusive effects are increased. Here, (a,b) $\varepsilon =1/Pe_s=5$, (c,d) $\varepsilon =1/Pe_s=0.5$, and (e, f) $\varepsilon =1/Pe_s=0.05$.

Mathematically, it can be seen from (7.6) and (7.7) that the effect of fast reaction and diffusion corresponds to shifting the poles of the solution further away from the real axis into the non-physical upper half-plane.

The bulk concentration and velocity fields can be found readily from the analytical solution. From (6.11),

(7.14)\begin{equation} \bar{q}(z) ={-}\frac{\mathrm{i}}{k}\,(h(z)+ \dot{\gamma}z), \end{equation}

which implies

(7.15)\begin{equation} q(z) = \frac{\mathrm{i}}{k} \left( 2\dot{\gamma} z + 2\varepsilon\kappa\varDelta-2\varepsilon\kappa\, \frac{\bar{U}(a-1,\kappa z + \varDelta)}{\bar{U}(a,\kappa z+ \varDelta)}\right), \end{equation}

where $\bar {U}$ denotes the Schwarz conjugate of the principal parabolic cylinder function. By (5.13), the bulk concentration field can therefore be determined by taking the real part of (7.15). Flow streamlines and contours of the concentration field are shown in figure 3 for varying values of $\delta$ and fixed $\varepsilon = 0.5$. Figures 2 and 3 demonstrate that steady remobilization occurs as both the surfactant solubility and its surface diffusivity are increased.

Figure 3.

Steady remobilization of the interface and its effect on the bulk concentration and velocity fields. Streamlines of the flow are shown in the lower half-plane, while bulk concentration field contours are shown in the upper half-plane. Interfacial surfactant concentrations $\varGamma (x)$ (in red) and slip velocities $\mathcal {U}(x)$ (in blue) are also displayed. Here, $l=\dot \gamma =1$, $\varepsilon =1/Pe_s=0.5$, and (a) $\delta =0.001$, (b) $\delta =0.1$, (c) $\delta =0.5$, and (d) $\delta =1$.

8. Steady equilibria at $Pe_s=\infty$

Surface diffusion might be expected naturally to mollify surfactant gradients on the interface and, as a result, to remobilize the interface, but the effect (on remobilization) of fast reactive exchange of surfactant between the bulk and the interface is less intuitive. The results of the previous section suggest that fast reactive exchange, at any value of the surface diffusion, assists with this remobilization. To explore this further, this section leverages the fact that the mathematical approach allows for any surface Péclet number to be considered. Consequently, the extreme case of no surface diffusion at all is studied theoretically. Such a dramatic difference in bulk diffusion and surface diffusion might not be expected physically, but since it can be studied in an explicit mathematical way, it is worth doing so on theoretical grounds. Moreover, in § 11, it is argued that even this apparently extreme scenario is not so far removed from physical reality.

In the limit of infinite surface Péclet number $Pe_s \to \infty$, the steady version of (6.12) reduces to

(8.1)\begin{equation} (h + {\rm i} \delta )\,\frac{\partial h }{\partial z} = \dot \gamma^2 z - {\rm i} \delta \dot \gamma.\end{equation}

This can be integrated directly with respect to $z$ to give

(8.2)\begin{equation} h^2 + 2{\rm i} \delta h = \dot \gamma^2 \left( z^2 - \frac{2 {\rm i} \delta }{\dot \gamma}\,z - l^2 \right), \end{equation}

with the same choice of integration constant as in the previous section. One advantage is immediately apparent: this is not a first-order ODE for $h(z)$, but an algebraic expression, and the parabolic cylinder equation is not relevant in this case. Expression (8.2) is actually just a quadratic equation for $h$ that is solved readily to give

(8.3)\begin{equation} h(z)={-} \dot \gamma \left( \frac{{\rm i} \delta }{\dot \gamma} + \left[ \left(z - \frac{{\rm i} \delta }{\dot \gamma} \right)^2 - l^2 \right]^{1/2} \right).\end{equation}

The real and imaginary parts of this analytic function, evaluated on $y=0$, give the equilibrium slip velocity and steady-state surfactant concentration. It is easy to see from (8.3) that $h(z)$ is lower analytic; indeed, its only singularities in the finite complex plane are square root branch points at the two locations

(8.4)\begin{equation} \pm l + \frac{{\rm i} \delta }{\dot \gamma}, \end{equation}

which, for $\dot \gamma, \delta > 0$, are strictly in the upper half-plane. It can also be verified that $\varGamma (x) \ge 0$ everywhere on the interface, which is another requirement for a physically admissible equilibrium surfactant concentration (Crowdy Reference Crowdy2021a,Reference Crowdyb).

From (6.11), the corresponding expression (Schwarz conjugate) for the complex potential for the bulk concentration is

(8.5)\begin{equation} \bar{q}(z) ={-}\frac{{\rm i} }{k}\,(h(z) +\dot \gamma z ) ={-}\frac{{\rm i} }{k} \left(\dot \gamma z -{\rm i} \delta - \dot \gamma \left[ \left(z - \frac{{\rm i} \delta }{\dot \gamma} \right)^2 - l^2 \right]^{1/2}\right),\end{equation}

whence

(8.6)\begin{equation} {q}(z) = \frac{{\rm i} \delta \dot \gamma }{\alpha} \left( z +\frac{{\rm i} \delta }{\dot \gamma} - \left[ \left(z + \frac{{\rm i} \delta }{\dot \gamma} \right)^2 - l^2 \right]^{1/2}\right), \end{equation}

where we have used (5.5) and (6.13). It can be checked that this function is analytic in the upper half-plane $y> 0$, decaying like $1/z$, as required. Indeed, this function has two square root branch points in the lower half, plane at

(8.7)\begin{equation} \pm l - \frac{{\rm i} \delta }{\dot \gamma}. \end{equation}

The steady-state bulk surfactant concentration $c(x,y)$ is given by

(8.8)\begin{equation} c(x,y)= {\rm Re} [ q(z) ], \quad y > 0. \end{equation}

It can be checked that this quantity is non-negative everywhere in the upper half-plane.

In the limit $\delta /\dot \gamma \to 0$, we have $c(x,y) \to 0$, and the two real points $\pm l$ are the edges of a branch cut of $h(z)$ on the interface; that is, there is an immobilized cap $[-l,l]$ occupied by surfactant with an elliptical concentration profile, where the slip velocity vanishes. There is no surfactant outside this cap. Indeed, in this limit, the steady stagnant cap solution found in Crowdy (Reference Crowdy2021a) for insoluble surfactant at infinite surface Péclet number is recovered.

For a fixed value of $l$, the key non-dimensional parameter determining the qualitative shape of the steady-state equilibria, and thus the degree of remobilization, is clearly

(8.9)\begin{equation} \frac{\delta }{\dot \gamma} = \frac{D c_0 }{\varGamma_\infty L \dot \gamma_0}\,\frac{1 }{k}. \end{equation}

It is interesting that in this dilute limit, this determining parameter is independent of $\beta$, the coefficient in the linear equation of state. Figure 4 shows streamlines in the lower half-plane, bulk concentration contours in the upper half-plane, and superposed graphs of $\mathcal {U}(x)$ and $\varGamma (x)$ for equilibria with $\delta /\dot \gamma = 0.01, 0.025, 0.1, 0.5$ and $l= \dot \gamma =1$. It is clear that the extent of the stagnant cap reduces as $\delta /\dot \gamma$ increases, the interface is remobilized, and the gradients of surfactant concentration decrease, leading to generally lower Marangoni stresses. These results can be compared with those in § 7 and, in particular, the results for $l=1$, $\delta /\dot {\gamma }=0.5$ when $Pe_s=\infty$ in figure 4(d), and for the same parameters but $Pe_s=2$ in figure 3(c). As expected, the remobilization is more enhanced when $Pe_s=2$, the gradients of $\varGamma (x)$ are smaller, and consequently the Marangoni stresses are also smaller. The analytical formulas for $Pe_s=\infty$ provide, in a sense, a worst-case scenario for remobilization in the fast kinetics exchange regime. The value of these $Pe_s = \infty$ solutions lies in their elegance and simplicity.

Figure 4.

Steady remobilization of the interface for increasing $\delta /\dot \gamma > 0$ for $Pe_s =\infty$. Streamlines (lower half-plane), bulk concentration contours (upper half-plane), and superposed graphs of $\mathcal {U}(x)$ (blue) and $\varGamma (x)$ (red) are shown for equilibria with $l= \dot \gamma =1$ and (a) $\delta /\dot \gamma = 0.01$, (b) $\delta /\dot \gamma = 0.025$, (c) $\delta /\dot \gamma = 0.1$, and (d) $\delta /\dot \gamma = 0.5$.

Mathematically, as indicated schematically in figure 5, the effect of the fast exchange kinetics is to shift up into the upper half-plane, by a distance $\delta /\dot \gamma$, a branch cut associated with the analytic function $h(z)$ determining the Stokes flow in the lower half-plane. This branch cut sits on the interface when $\delta /\dot \gamma =0$, and corresponds to the stagnant cap, but fast reaction with the upper gas phase, or $\delta /\dot \gamma > 0$, shifts it off the interface into the non-physical upper half-plane, thereby remobilizing the stagnant cap.

Figure 5.

Fast reaction moves the branch cut of $h(z)$ up by distance $\delta /\dot \gamma$ into the non-physical upper half-plane. (a) When $\delta /\dot \gamma =0$, the branch cut corresponds to a weak solution (stagnant cap on the interface). (b) For $\delta /\dot \gamma >0$, the branch cut moves off the interface, smoothing out the solution and remobilizing the interface.

9. General implicit solution for the unsteady evolution

This complex variable reformulation of the problem allows the analysis for $Pe_s =\infty$ to be carried much further. Indeed, it is now shown that an implicit form of the general solution for the time-dependent problem can be found using a complex generalization of the method of characteristics introduced for Marangoni flows of this kind by Crowdy (Reference Crowdy2021a,Reference Crowdyb). This allows us to capture analytically the unsteady evolution to the steady equilibria found in § 8. The following solutions are generalizations – to account for fast reactive surfactant exchange with the upper fluid – of the exact solutions for the formation of stagnant caps found in Crowdy (Reference Crowdy2021a).

First, introduce $g(z,t)$ with

(9.1)\begin{equation} h(z,t) ={-}\dot \gamma z + g(z,t),\end{equation}

where we assume that

(9.2)\begin{equation} g(z,t) = {{O}}(1/z) \quad {\rm as}\ |z| \to \infty,\end{equation}

which is consistent with the assumption that the far-field interface is clean. From (6.11) and (9.1),

(9.3)\begin{equation} \bar{q}(z,t) ={-}\frac{{\rm i} }{k}\,g(z,t),\end{equation}

so that the complex potential for the bulk concentration in $y > 0$ is

(9.4)\begin{equation} {q}(z,t) = \frac{{\rm i} }{k}\,\bar{g}(z,t).\end{equation}

Equation (6.12) with $Pe_s \to \infty$ becomes

(9.5)\begin{equation} \frac{\partial h }{\partial t} + h\,\frac{\partial h }{\partial z} + {\rm i} \delta \left( \frac{\partial h }{\partial z} + \dot \gamma \right) =\dot \gamma^2 z, \end{equation}

which, on substitution of (9.1), is

(9.6) \begin{equation} \frac{\partial g }{\partial t} + (-\dot \gamma z + g) \left(-\dot \gamma + \frac{\partial g }{\partial z} \right) + {\rm i} \delta\,\frac{\partial g}{\partial z} = \dot \gamma^2 z, \end{equation}

which simplifies to

(9.7)\begin{equation} \frac{\partial g }{\partial t} + (-\dot \gamma z + g + {\rm i} \delta )\, \frac{\partial g }{\partial z} = \dot \gamma g.\end{equation}

This can be solved by a complex method of characteristics as discussed in Crowdy (Reference Crowdy2021a,Reference Crowdyb) for a complex Burgers type PDE similar to (9.7). In this case, this is done by restating it as

(9.8)\begin{equation} \frac{{\rm d} g }{{\rm d} t} = \dot \gamma g, \quad g(z,0)=G(Z),\end{equation}

on complex characteristics defined by

(9.9)\begin{equation} \frac{{\rm d} z }{{\rm d} t} ={-}\dot \gamma z + g(z,t) + {\rm i} \delta, \quad z=Z, \quad {\rm at}\ t=0.\end{equation}

The parameter $Z$ can be thought of as a label of characteristics. Solving (9.8) by an integrating factor leads to

(9.10)\begin{equation} {\rm e}^{-\dot \gamma t}\,g(z,t) = G(Z). \end{equation}

Use of this in (9.9) leads to

(9.11)\begin{equation} \frac{{\rm d} z }{{\rm d} t} + \dot \gamma z = {\rm e}^{\dot \gamma t}\,G(Z) + {\rm i} \delta, \quad z=Z, \quad {\rm at}\ t=0. \end{equation}

Multiplication by the integrating factor ${\rm e}^{\dot \gamma t}$ turns this into

(9.12)\begin{equation} \frac{{\rm d} ({\rm e}^{\dot \gamma t} z) }{{\rm d} t} = {\rm e}^{2\dot \gamma t}\,G(Z) + {\rm i} \delta \,{\rm e}^{\dot \gamma t}, \quad z=Z, \quad {\rm at} \ t=0, \end{equation}

which, on integration, leads to

(9.13)\begin{equation} {\rm e}^{\dot \gamma t}\,z = r_2(t)\,G(Z) + {\rm i}\delta\,r_1(t) + Z, \end{equation}

where

(9.14a,b)\begin{equation} r_1(t) = \frac{{\rm e}^{\dot \gamma t} - 1 }{\dot \gamma}, \quad r_2(t) = \frac{{\rm e}^{2\dot \gamma t} -1 }{2\dot \gamma}, \end{equation}

which have the property that $r_1(0)=r_2(0)=0$.

In summary, we have found an implicit form of the general solution to be

(9.15)\begin{equation} g(z,t) = G(Z) \,{\rm e}^{\dot \gamma t},\end{equation}

where

(9.16)\begin{equation} z = {\rm e}^{-\dot \gamma t}\,r_2(t)\,G(Z) + {\rm i}\delta \,{\rm e}^{-\dot \gamma t}\,r_1(t) + Z \,{\rm e}^{-\dot \gamma t}. \end{equation}

To solve a given problem, the initial data values for $G(Z)=g(Z,0)$ need to be provided. These initial data need to satisfy certain constraints and cannot be chosen arbitrarily (Crowdy Reference Crowdy2021a,Reference Crowdyb).

10. A class of explicit time-evolving solutions

Remarkably, as found for closely related problems by Crowdy (Reference Crowdy2021a,Reference Crowdyb), a particular choice of initial conditions leads to fully explicit time-evolving solutions. Let initial conditions be given by

(10.1)\begin{equation} G(Z) = \frac{A }{Z - {\rm i} B}, \quad A, B \in \mathbb{R}, \quad A, B>0,\end{equation}

which is lower analytic and satisfies the requirement (9.2). The initial surfactant distribution $\varGamma (x,0)$ and slip velocity ${\mathcal {U}}(x,0)$ associated with (10.1) are found easily to be

(10.2)\begin{equation} \varGamma(x,0) =\frac{AB }{x^2+B^2}, \quad {\mathcal{U}}(x,0) ={-} \dot \gamma x + \frac{A x}{x^2+B^2}. \end{equation}

The initial condition (10.1) has been chosen carefully: the parameter $B$ must be positive to ensure that $g(z,0)$ is lower analytic, while $A$ must be positive to ensure that the bulk and surface concentrations are non-negative quantities; Crowdy (Reference Crowdy2021a,Reference Crowdyb) discusses such matters in more detail in the context of similar Marangoni flow problems. Use of (10.1) in (9.16) leads to

(10.3)\begin{equation} {\rm e}^{\dot \gamma t} z = \frac{A\,r_2(t) }{Z- {\rm i} B} + {\rm i}\delta\,r_1(t) + Z.\end{equation}

Significantly, this can be rearranged to form the quadratic equation

(10.4)\begin{equation} Z^2 - ({\rm i} B - {\rm i} \delta\,r_1(t) + {\rm e}^{\dot \gamma t}\,z) Z + A\,r_2(t) +{\rm i} B(-{\rm i} \delta\,r_1(t) + {\rm e}^{\dot \gamma t}\,z) = 0. \end{equation}

This has the solution, satisfying the initial condition $Z=z$ at $t=0$, given by

(10.5)\begin{align} Z = Z(z,t) &= \frac{1 }{2} \left({\rm i} B - {\rm i} \delta\,r_1(t) + {\rm e}^{\dot \gamma t}\,z \vphantom{\frac{A r_2(t) }{Z- {\rm i} B}} \right.\nonumber\\ &\quad - ({\rm i} B + {\rm i} \delta\,r_1(t) - {\rm e}^{\dot \gamma t}\,z) \left.\left[1 - \frac{4 A\, r_2(t) }{({\rm i} B + {\rm i} \delta\,r_1(t) - {\rm e}^{\dot \gamma t}\,z)^2}\right]^{1/2} \right). \end{align}

The solution for $g(z,t)$ is then given explicitly as

(10.6)\begin{equation} g(z,t) = \frac{A\, {\rm e}^{\dot \gamma t}}{Z(z,t) -{\rm i} B}, \end{equation}

so that the corresponding $h(z,t)$ and $\bar {q}(z,t)$, following respectively from (9.1) and (9.3), are

(10.7a,b)\begin{equation} h(z,t) ={-}\dot \gamma z + \frac{A \,{\rm e}^{\dot \gamma t} }{Z(z,t)-{\rm i} B}, \quad \bar{q}(z,t) ={-}\frac{{\rm i} \kappa A \,{\rm e}^{\dot \gamma t} }{Z(z,t)-{\rm i} B}. \end{equation}

It follows that the complex potential describing the bulk concentration is given by

(10.8)\begin{equation} {q}(z,t) = \frac{{\rm i} \kappa A \,{\rm e}^{\dot \gamma t} }{\bar{Z}(z,t)+{\rm i} B}, \end{equation}

where

(10.9)\begin{align} \bar{Z}(z,t) &= \frac{1 }{2} \left(-{\rm i} B + {\rm i} \delta\,r_1(t) + {\rm e}^{\dot \gamma t}\,z \vphantom{\frac{A r_2(t) }{Z- {\rm i} B}} \right.\nonumber\\ &\quad +({\rm i} B + {\rm i} \delta\,r_1(t) + {\rm e}^{\dot \gamma t}\,z) \left.\left[1 - \frac{4 A\,r_2(t) }{({\rm i} B + {\rm i} \delta\,r_1(t) + {\rm e}^{\dot \gamma t}\,z)^2}\right]^{1/2} \right). \end{align}

It is of interest to examine the large-time asymptotics of this solution class. It is straightforward to show from (10.5) that as $t \to \infty$,

(10.10)\begin{equation} {\rm e}^{-\dot \gamma t}\,Z \to \frac{1 }{2} \left[ \left(z-\frac{{\rm i} \delta }{\dot \gamma} \right) + \left( \left(z-\frac{{\rm i} \delta }{\dot \gamma} \right)^2 -\frac{2 A }{\dot \gamma} \right)^{1/2} \right], \end{equation}

and from (10.7a,b), that as $t \to \infty$,

(10.11)\begin{equation} h(z,t) \to -\dot \gamma z + \frac{A }{{\rm e}^{-\dot \gamma t}\,Z(z,t)}. \end{equation}

On combining (10.10) and (10.11), it can be shown further that as $t \to \infty$,

(10.12)\begin{equation} h(z,t) \to - {\rm i} \delta - \dot \gamma \left( \left(z-\frac{{\rm i} \delta }{\dot \gamma} \right)^2 -\frac{2 A }{\dot \gamma} \right)^{1/2}. \end{equation}

The solution therefore tends at large times to one of the steady equilibria (8.3) found in § 8, with the parameter $l$ determined by initial conditions according to

(10.13)\begin{equation} l = \sqrt\frac{2A }{\dot \gamma}. \end{equation}

To illustrate the unsteady dynamics, figure 6 shows a time sequence (snapshots) of the instantaneous streamlines, bulk concentration contours, and graphs of $\mathcal {U}(x,t) = {\rm Re} [h(x,t)]$ and $\varGamma (x,t) = {\rm Im} [h(x,t)]$ for an example calculation with parameter choices $A=0.5$, $B=0.25$, $\delta = 0.05$ and $\dot \gamma =1$. By $t=1$, the dynamical evolution has drawn close to the equilibrium solution. An interesting feature is that Marangoni stresses acting against those from the ambient strain are initially sufficiently large to cause a pair of recirculating viscous eddies in the fluid; these eventually die out as the steady equilibrium is reached. In this example, $\delta /\dot \gamma = 0.05$ is sufficiently small that the final equilibrium is only weakly remobilized and is still close to the stagnant cap equilibrium. Figure 7 shows another series of snapshots for $A=0.5$, $B=0.25$, $\delta = 0.5$ and $\dot \gamma =1$ when the degree of eventual remobilization at equilibrium is more pronounced. By $t=1.2$, the dynamical evolution has drawn close to the equilibrium solution.

Figure 6.

Unsteady evolution to a weakly remobilized steady state. Time snapshots at (a) $t=0$, (b) $t=0.2$, (c) $t=0.4$, (d) $t=0.6$, (e) $t=0.8$, and ( f) $t=1$. Instantaneous streamlines, bulk concentration contours, and graphs of $\mathcal {U}(x,t)$ and $\varGamma (x,t)$ are shown for $A=0.5$, $B=0.25$, $\delta = 0.05$ and $\dot \gamma =1$.

Figure 7.

Unsteady evolution to a more strongly remobilized steady state. Time snapshots at (a) $t=0$, (b) $t=0.1$, (c) $t=0.2$, (d) $t=0.4$, (e) $t=0.8$, and ( f) $t=1.2$. Instantaneous streamlines, bulk concentration contours, and graphs of $\mathcal {U}(x,t)$ and $\varGamma (x,t)$ are shown for $A=0.5$, $B=0.25$, $\delta = 0.5$ and $\dot \gamma =1$.

Other choices of initial condition are found to give qualitatively similar dynamics, with the nature of the final remobilized interface governed by the value of $\delta /\dot \gamma$.

11. Physical systems and justification of the mathematical model

Wang et al. (Reference Wang, Papageorgiou and Maldarelli1999,  Reference Wang, Papageorgiou and Maldarelli2002) justify their diffusion-limited study of rising bubbles by providing several examples where $Bi$ is large – see Table 1 in Wang et al. (Reference Wang, Papageorgiou and Maldarelli2002). The surfactants used are three to seven chain length alcohols (hexanol), and common polyethylene oxide ones such as ${\rm C}_{10}{\rm E}_8$ and ${\rm C}_{12}{\rm E}_8$. Using values $R=8.3\ {\rm J\ K}^{-1}\ {\rm mol}^{-1}$ and $T=273\ {\rm K}$ gives $\beta \approx 2.3\times 10^3\ {\rm J}\ {\rm mol}^{-1}$. The maximum packing concentration for hexanol was measured by Joos & Serrien (Reference Joos and Serrien1989) to be $\varGamma _\infty \approx 3\times 10^{-6}\ {\rm mol}\ {\rm m}^{-2}$, and Chang & Franses (Reference Chang and Franses1995) give values $2\unicode{x2013}6\times 10^{-6}\ {\rm mol}\ {\rm m}^{-2}$. The kinetic rate constants are harder to establish, but Chang & Franses (Reference Chang and Franses1995) give a range of values for the desorption constant $a$ to be $10^{-4}\unicode{x2013}10^2\ {\rm s}^{-1}$. For glycerol mixtures, we have viscosities $\mu \approx 1.5\ {\rm kg}\ {\rm m}^{-1}\ {\rm s}^{-1}$, so the velocity scale $\mathcal {U}_0$ given by (3.1) is in the range $\mathcal {U}_0\approx 10^{-3}\ {\rm m}\ {\rm s}^{-1}$. As mentioned in § 3, the velocity scale $\mathcal {U}_0^*=\dot {\gamma }_0 L$ is more representative in estimating crucial parameters such as the Biot and Péclet numbers. The velocity $\mathcal {U}_0$ is naturally associated with the spreading of surfactant at an interface between two immiscible fluids with viscosities $\mu _1$ and $\mu _2$.

Considering the results of figures 3 and 4, it is clear that remobilization takes place when the parameter $\delta /\dot {\gamma }$ increases from $0.01$ to approximately $0.5$ or larger. Since $\dot {\gamma }=\dot {\gamma }_0/(\mathcal {U}_0/L)$, and noting that $\mathcal {U}_0/L$ is fixed, to remobilize, the strain rate $\dot {\gamma }_0$ must decrease, i.e. $\dot {\gamma }$ decreases by a factor 50 in going from $\delta /\dot {\gamma }=0.01$ to $0.5$. Defining Biot and Péclet numbers using $\mathcal {U}_0^*$, namely $Bi=aL/\mathcal {U}_0^*$, $Pe=\mathcal {U}_0^* L/D$, shows immediately that $Bi$ increases by a factor of $50$ while $Pe$ decreases by the same factor. An estimate can be found by considering $L$ to be $1\unicode{x2013}10\ \mathrm {\mu }{\rm m}$, say, while the diffusion coefficient is $D=10^{-10}\ {\rm m}^2\ {\rm s}^{-1}$ (Chang & Franses Reference Chang and Franses1995). Taking the typical velocity when $\delta /\dot {\gamma }=0.01$ to be $\mathcal {U}_0\approx 10^{-3}\ {\rm m}\ {\rm s}^{-1}$, the Biot and Péclet numbers can be estimated at $\delta /\dot {\gamma }=0.5$ to be $Bi\approx 2.5a$ and $Pe\approx 1$. Using values for the desorption constant $a$ (see above), we have $2.5\times 10^{-4}\lessapprox Bi\lessapprox 2.5\times 10^2$, which places us within the reach of the theory when the desorption constant is not too small.