4.1. Static meniscus

At rest, the velocity field $\boldsymbol {u}_0$ is null everywhere and the pressure is hydrostatic, i.e. $p_0=-z$. Therefore, the static configuration is obtained by solving the nonlinear equation associated with the shape of the axisymmetric static meniscus, $\eta _0(r)$,

(4.1)\begin{equation} \eta_0-\frac{\kappa\left(\eta_0\right)}{Bo}=0, \end{equation}

with $\kappa (\eta _0)=(\eta _{0,rr}+\eta _{0,r}(1+\eta _{0,r}^2)/r)(1+\eta _{0,r}^2)^{-3/2}$. At the centreline, $r=0$, the regularity condition $\eta _{0,r}=0$ holds due to axisymmetry. The shape of the meniscus is obtained by imposing the geometric relation at the contact line, $r=1$,

(4.2)\begin{equation} \left.\frac{\partial\eta_0}{\partial r}\right|_{r=1}=\cot{\theta_s}, \end{equation}

where $\theta _s$ is a prescribed static contact angle (see also figure 1a). When $\theta _s$ is set to ${\rm \pi} /2$, then the static interface appears flat.

4.2. Linear eigenvalue problem

Governing equations (2.1) and their boundary conditions (2.2) are then linearized around the static base-flow. It follows that at order $\epsilon$ the velocity and pressure fields satisfy the Stokes equations

(4.3)\begin{equation} \boldsymbol{\nabla}\boldsymbol{\cdot}\boldsymbol{u}_1=0,\quad \frac{\partial \boldsymbol{u}_1}{\partial t}+\boldsymbol{\nabla} p_1-\frac{1}{Re}\Delta\boldsymbol{u}_1=\boldsymbol{0}, \end{equation}

with the linearized kinematic and dynamic free surface boundary conditions (at $z=\eta _0$)

(4.4)\begin{gather} \frac{\partial\eta_1}{\partial t}+{\left.u_{r,1}\right|_{\eta_0}\frac{\partial\eta_0}{\partial r}-\left.u_{z,1}\right|_{\eta_0}}=0, \end{gather}

(4.5)\begin{gather} -\left.p_1\right|_{\eta_0}\boldsymbol{n}\left(\eta_0\right)+\eta_1\boldsymbol{n}\left(\eta_0\right)+\frac{1}{Re}\left.\left(\boldsymbol{\nabla}\boldsymbol{u}_1+\boldsymbol{\nabla}^{\rm T}\boldsymbol{u}_1\right)\right|_{\eta_0}\boldsymbol{\cdot}\boldsymbol{n}\left(\eta_0\right)=\frac{1}{Bo}\left.\frac{\partial\kappa\left(\eta\right)}{\partial\eta}\right|_{\eta_0}\eta_1\boldsymbol{n}\left(\eta_0\right), \end{gather}

where $\boldsymbol {n}(\eta _0)=\{-\eta _{0,r},0,1\}^{\rm T}(1+\eta _{0,r}^2)^{-1/2}$ and

(4.6)\begin{align} \left.\frac{\partial\kappa\left(\eta\right)}{\partial\eta}\right|_{\eta_0}\eta_1 &= \frac{\left(1+\eta_{0,r}^2\right)-3r\eta_{0,r}\eta_{0,rr}}{\left(1+\eta_{0,r}^2\right)^{5/2}}\frac{1}{r}\frac{\partial\eta_1}{\partial r}+\frac{1}{\left(1+\eta_{0,r}^2\right)^{3/2}}\frac{\partial^2\eta_1}{\partial r^2}\nonumber\\ &\quad +\frac{1}{\left(1+\eta_{0,r}^2\right)^{1/2}}\frac{1}{r^2}\frac{\partial^2\eta_1}{\partial\phi^2} \end{align}

is the first-order variation of the curvature associated with the small perturbation $\epsilon \eta _1$. The azimuthal coordinate is denoted by $\phi$. The no-slip boundary condition is imposed at the solid walls, $\boldsymbol {u}_1=\boldsymbol {0}$, while the pinned contact line condition is enforced at the contact line, $z=\eta _0$ and $r=1$,

(4.7)\begin{equation} \left.\frac{\partial\eta_1}{\partial t}\right|_{r=1}=0. \end{equation}

Hence, the linear system can be written in compact form as

(4.8)\begin{equation} \left({{\boldsymbol{\mathsf{B}}}}\partial_t-{{\boldsymbol{\mathsf{A}}}}\right)\boldsymbol{q}_1=\boldsymbol{0},\quad \text{with}\ {{\boldsymbol{\mathsf{B}}}}=\left( \begin{matrix} I & 0\\ 0 & 0 \end{matrix} \right),\quad {{\boldsymbol{\mathsf{A}}}}=\left( \begin{matrix} Re^{{-}1}\Delta & -\boldsymbol{\nabla}\\ \boldsymbol{\nabla}^{\rm T} & 0 \end{matrix} \right). \end{equation}

We note that the kinematic and the dynamic boundary conditions (4.4) and (4.5) do not explicitly appear in (4.8), but they are imposed as conditions at the interface (Viola & Gallaire Reference Viola and Gallaire2018). More precisely, in the numerical scheme, the kinematic condition governing the state variable $\eta$ is implemented as an additional equation dynamically coupled with $\boldsymbol {u}_1$ and $p_1$ in (4.8) (this is better clarified in Appendix B of Bongarzone, Viola & Gallaire (Reference Bongarzone, Viola and Gallaire2021b)), whereas the three stress components of the dynamic condition are enforced as standard boundary conditions in the corresponding components of the momentum equation. The solution can be then expanded in terms of normal modes in time and in the azimuthal direction as

(4.9)\begin{align} \boldsymbol{q}_1\left(r,\phi,z,t\right)&=\hat{\boldsymbol{q}}_1\left(r,z\right){\rm e}^{\lambda t}{\rm e}^{\text{i}m\phi}+{\rm c.c.}\nonumber\\ &\quad=\left\{\hat{u}_{1r}\left(r,z\right),\hat{u}_{1\phi}\left(r,z\right),\hat{u}_{1z}\left(r,z\right),\hat{p}_1\left(r,z\right)\right\}^{\rm T} {\rm e}^{\lambda t}{\rm e}^{\text{i}m\phi}+{\rm c.c.}, \end{align}

(4.10)\begin{align} &\qquad \qquad\eta_1\left(r,\phi,t\right)=\hat{\eta}_1\left(r\right){\rm e}^{\lambda t}{\rm e}^{\text{i}m\phi}+{\rm c.c.} \end{align}

Substituting the normal form (4.9)–(4.10) in system (4.8), we obtain a generalized linear eigenvalue problem,

(4.11)\begin{equation} \left(\lambda{{\boldsymbol{\mathsf{B}}}}-{{\boldsymbol{\mathsf{A}}}}_m\right)\hat{\boldsymbol{q}}_1=\boldsymbol{0}, \end{equation}

where the linear operator ${{\boldsymbol{\mathsf{A}}}}_m$ depends on the azimuthal wavenumber $m$ and $\hat {\boldsymbol {q}}_1$ is the global mode associated with the eigenvalue $\lambda =-\sigma +\text {i}\omega$, with $\sigma$ and $\omega$ the damping coefficient and the natural frequency, respectively, of the $(m,n)$ global mode. Here the indices $(m,n)$ represent the number of nodal circles and nodal diameters, respectively. Owing to the normal mode expansion (4.9), we notice that the operator ${{\boldsymbol{\mathsf{A}}}}_m$ is complex, since $\phi$ derivatives produce $\text {i}m$ terms. A complete expansion of the complex operator can be found in Meliga, Chomaz & Sipp (Reference Meliga, Chomaz and Sipp2009) and Viola & Gallaire (Reference Viola and Gallaire2018).

In order to regularize the problem at the axis, depending on the selected azimuthal wavenumber $m$, different regularity conditions must be imposed at $r=0$ (Liu & Liu Reference Liu and Liu2012; Viola & Gallaire Reference Viola and Gallaire2018),

(4.12a)\begin{gather} |m|=0\text{:}\quad \hat{u}_{1r}=\hat{u}_{1\phi}=\frac{\partial\hat{u}_{1z}}{\partial r}=\frac{\partial \hat{p}_1}{\partial r}=0, \end{gather}

(4.12b)\begin{gather} |m|=1\text{:}\quad \frac{\partial\hat{u}_{1r}}{\partial r}=\frac{\partial\hat{u}_{1\phi}}{\partial r}=\hat{u}_{1z}=\hat{p}_1=0, \end{gather}

(4.12c)\begin{gather} |m|>0\text{:}\quad \hat{u}_{1r}=\hat{u}_{1\phi}=\hat{u}_{1z}=\hat{p}_1=0. \end{gather}

Lastly, due to the symmetries of the problem, system (4.11) with its boundary conditions is invariant under the

(4.13)\begin{equation} \left(\hat{u}_{1r},\hat{u}_{1\phi},\hat{u}_{1z},\hat{p}_1,\hat{\eta}_1,+m,-\sigma+\text{i}\omega\right)\rightarrow\left(\hat{u}_{1r},-\hat{u}_{1\phi},\hat{u}_{1z},\hat{p}_1,\hat{\eta}_1,-m,-\sigma+\text{i}\omega\right), \end{equation}

transformation, so in this section, § 4, we consider only the case with $m\geqslant 0$. Furthermore, the following relations hold:

(4.14)\begin{gather} \left(\hat{\boldsymbol{q}}_1,\hat{\eta}_1,+m,-\sigma+\text{i}\omega\right)\rightarrow\left(\hat{\boldsymbol{q}}_1^*,\hat{\eta}_1^*,-m,-\sigma-\text{i}\omega\right), \end{gather}

(4.15)\begin{gather} \left(\hat{\boldsymbol{q}}_1,\hat{\eta}_1,-m,-\sigma+\text{i}\omega\right)\rightarrow\left(\hat{\boldsymbol{q}}_1^*,\hat{\eta}_1^*,+m,-\sigma-\text{i}\omega\right), \end{gather}

(where the star designates the complex conjugate), i.e. the eigenvalues are complex conjugates and all spectra ($\pm m$) in the $(\sigma,\omega )$-plane are symmetric with respect to the real axis ($\omega =0$), but the complex conjugates of the corresponding eigenvectors, except for the axisymmetric dynamics ($m=0$), are not eigenmodes of the same spectrum. The damping coefficients and natural frequencies of viscous capillary–gravity waves with fixed contact line in both the meniscus-free and with-meniscus configuration are thus computed by solving numerically the generalized eigenvalue problem (4.11), as described in § 3.

With regard to the literature survey outlined in table 1, in Appendix A we propose a thorough validation of our numerical tools via comparison with several pre-existing experiments and theoretical/semianalytical predictions focusing on both brimful and nearly brimful circular cylinders.

Table 1.

Literature survey on the natural frequencies and damping coefficients of small-amplitude capillary–gravity waves in labscale upright cylindrical containers with pinned contact line and in both meniscus-free and with-meniscus configurations. The present work lies within the conditions highlighted by the shaded frames. The case examined by K13 and S21 will be discussed afterwards in § 5 within the context of subharmonic Faraday waves.

5.1. Presentation

Here, the full system (2.1)–(2.3) is solved through a WNL analysis based on the multiple scale method that is valid in the regime of small perturbations of the static configuration and small external control parameters, namely the driving amplitude and detuning from the parametric resonance. Let us thus introduce the following asymptotic expansion for the flow quantities:

(5.1)\begin{gather} \boldsymbol{q}=\left\{\boldsymbol{u},p\right\}^{\rm T}=\boldsymbol{q}_0+\epsilon\boldsymbol{q}_1+\epsilon^2\boldsymbol{q}_2+\epsilon^3\boldsymbol{q}_3+O\left(\epsilon^4\right), \end{gather}

(5.2)\begin{gather} \eta=\eta_0+\epsilon\eta_1+\epsilon^2\eta_2+\epsilon^3\eta_3+O\left(\epsilon^4\right). \end{gather}

In the spirit of the multiple scale technique, we introduce the slow time scale $T=\epsilon ^2 t$, with $t$ being the fast time scale at which the free surface oscillates. For subharmonic resonances the system is expected to respond with a frequency equal to half the driving frequency. In order to determine the boundaries of the instability tongues, we assume the external forcing angular frequency to be $\varOmega _d=2\omega +\varLambda$, where $\omega$ is the natural frequency associated with the generic $(m,n)$ capillary–gravity wave considered and $\varLambda$ is the detuning parameter. As, by construction, the WNL analysis is valid close to the instability threshold only, we assume a departure from criticality to be of order $\epsilon ^2$. In terms of control parameters, this assumption translates to the following scalings for the external forcing amplitude, $F_d/g$, and detuning $\varLambda$:

(5.3)\begin{equation} F_d/g=F=\epsilon^2\hat{F},\quad \varLambda=\epsilon^2\hat{\varLambda}. \end{equation}

It should be noted that the presence of viscosity leads to a damped $\epsilon$-order solution $\boldsymbol {q}_1$ (as discussed in § 4), whereas standard multiple scale methods apply to marginally stable systems (Nayfeh Reference Nayfeh2008). Nevertheless, as the Reynolds number is typically high enough, the damping coefficient results in a slow damping process over fast wave oscillations (see § 4). In such a regime, a multiple scale analysis can still be applied by postulating that the damping coefficient of the $(m,n)$ wave is of order $\epsilon ^2$, i.e. $\sigma =\epsilon ^2\hat {\sigma }$, therefore the $(m,n)$ eigenvalue reads $\lambda =-\epsilon ^2\hat {\sigma }+\text {i}\omega$. A simple way to account for this second-order departure from neutrality consists in replacing the leading-order operator ${{\boldsymbol{\mathsf{A}}}}_m={{\boldsymbol{\mathsf{A}}}}_m(Re)$ defined in (4.11), for which $\hat {\boldsymbol {q}}_1$ is not neutral, but rather stable, by the shifted operator (Meliga et al. Reference Meliga, Chomaz and Sipp2009), $\tilde {{{\boldsymbol{\mathsf{A}}}}}_m={{\boldsymbol{\mathsf{A}}}}_m+\epsilon ^2{{\boldsymbol{\mathsf{S}}}}_m$, where ${{\boldsymbol{\mathsf{S}}}}_m$ is the shift operator defined as ${{\boldsymbol{\mathsf{S}}}}_m\hat {\boldsymbol {q}}_1=-\hat {\sigma }\hat {\boldsymbol {q}}_1$. The shifted operator $\tilde {{{\boldsymbol{\mathsf{A}}}}}_m$ is characterized by the same spectra of ${{\boldsymbol{\mathsf{A}}}}_m$, except that the $(m,n)$ eigenmode $\hat {\boldsymbol {q}}_1$ associated with $\hat {\sigma }$ is now marginally stable, and hence the WNL formalism can be applied. For a thorough discussion about the formalism of the shift operator see Meliga et al. (Reference Meliga, Chomaz and Sipp2009) and Meliga, Gallaire & Chomaz (Reference Meliga, Gallaire and Chomaz2012). Although a different approach to account for a damped first-order solution was followed by Viola & Gallaire (Reference Viola and Gallaire2018), leading to a different (but equivalent) asymptotic expansion, we use in this paper the shift operator approach.

Finally, substituting the asymptotic expansions and scalings above in the governing equations (2.1)–(2.3) with their boundary conditions, a series of problems at the different orders in $\epsilon$ are obtained.

As anticipated in § 3, when contact angle effects are included in the analysis, i.e. the initial static interface is not flat, the third-order asymptotic expansion of the full viscous hydrodynamic system introduced in § 2 turns out to be very complex to derive analytically. Particularly tedious is the dynamic boundary condition, as it involves free surface boundary terms, which, within the linearization process, must be flattened at the static interface, $\eta _0$, as well as the full nonlinear curvature. In order to overcome these practical difficulties, the linearization and expansion procedures have been fully automated using symbolic calculus within the software Wolfram Mathematica, which has been then integrated within the main code implemented in MATLAB. The corresponding Mathematica codes are provided as supplementary material.

5.2. Order $\epsilon ^0$: static meniscus

At order $\epsilon ^0$ the system reduces to the nonlinear equation associated with the shape of the axisymmetric static meniscus. The velocity field is null, $\boldsymbol {u}_0=\boldsymbol {0}$ and the pressure is hydrostatic, $p_0=-z$. As described in § 4.1, the static interface, $\eta _0(r)$, is obtained by prescribing a static contact angle, $\theta _s$, which enters through the geometrical relation (4.2) imposed at the contact line.

5.3. Order $\epsilon$: capillary–gravity waves

At leading order in $\epsilon$ the system is represented by the unsteady Stokes equations (4.3), together with the kinematic and dynamic boundary conditions (4.4)–(4.5), linearized around the static base flow $\boldsymbol {q}_0=\{\boldsymbol {u}_0,p_0\}^{\rm T}=\{\boldsymbol {0},-z\}^{\rm T}$ and $\eta _0$, and subjected to the no-slip boundary condition at the solid walls, regularity conditions at the axis (4.12a)–(4.12c) and to the pinned contact line condition (4.7),

(5.4)\begin{equation} \left({{\boldsymbol{\mathsf{B}}}}\partial_t-\tilde{{{\boldsymbol{\mathsf{A}}}}}_m\right)\boldsymbol{q}_1=\boldsymbol{0}. \end{equation}

Within the framework of the Faraday instability, we are interested in a standing waveform of the solution, which can be seen as a result of the balance of two counter-rotating waves. Hence, we look for a first-order solution of the form

(5.5)$$\begin{gather} \boldsymbol{q}_1=A_1^+\left(T\right)\hat{\boldsymbol{q}}_1^{A^+} \exp\left({\text{i}\left(\omega t+m\phi\right)}\right)+A_1^-\left(T\right)\hat{\boldsymbol{q}}_1^{A^-} \exp\left({\text{i}\left(\omega t-m\phi\right)}\right)+{\rm c.c.}, \end{gather}$$

(5.6)$$\begin{gather}{\eta_1=A_1^+\left(T\right)\hat{\eta}_1^{A^+} \exp\left({\text{i}\left(\omega t+m\phi\right)}\right)+A_1^-\left(T\right)\hat{\eta}_1^{A^-} \exp\left({\text{i}\left(\omega t-m\phi\right)}\right)+{\rm c.c.},} \end{gather}$$

($\eta _1$ takes the same form) that destabilizes the static configuration. A single azimuthal wavenumber $m$ is considered at a time. In (5.5), $A^+_1$ and $A^-_1$, unknown at this stage of the expansion, are the complex amplitudes of the oscillating mode $\hat {\boldsymbol {q}}_1^{A^+}$ and $\hat {\boldsymbol {q}}_1^{A^-}$, respectively, and they are functions of the slow time scale $T$. The eigensolution of (5.4) has been widely discussed in § 4 for $m\geqslant 0$. We note in addition that the eigenmode for the $-m$ perturbation is similar to that of the $+m$ perturbation; more precisely, it oscillates with the same frequency $\omega$, but it has the opposite pitch and it rotates in the opposite direction.

5.4. Order $\epsilon ^2$: MW, second harmonics and mean-flow corrections

At order $\epsilon ^2$ we obtain the linearized Stokes equations and boundary conditions applied to $\boldsymbol {q}_2=\{\boldsymbol {u}_2,p_2\}^{\rm T}$ and $\eta _2$,

(5.7)\begin{equation} \left({{\boldsymbol{\mathsf{B}}}}\partial_t-\tilde{{{\boldsymbol{\mathsf{A}}}}}_m\right)\boldsymbol{q}_2=\boldsymbol{\mathcal{F}}_2, \end{equation}

and forced by a term $\boldsymbol {\mathcal {F}}_2$ depending only on zero-, first-order solutions and on the external forcing

(5.8)$$\begin{gather} \boldsymbol{\mathcal{F}}_2=|A^+_1|^2\hat{\boldsymbol{\mathcal{F}}}_2^{A^+A^{+^*}}+|A^-_1|^2\hat{\boldsymbol{\mathcal{F}}}_2^{A^-A^{-^*}}+\left(\hat{F}\hat{\boldsymbol{\mathcal{F}}}_2^{\hat{F}} \exp\left({\text{i}\left(2\omega t+\hat{\varLambda}T\right)}\right)+{\rm c.c.}\right)\nonumber\\ +\left(A^{+^2}_1\hat{\boldsymbol{\mathcal{F}}}_2^{A^+A^+} \exp\left({\text{i}\left(2\omega t+2m\phi\right)}\right)+A^{-^2}_1\hat{\boldsymbol{\mathcal{F}}}_2^{A^-A^-} \exp\left({\text{i}\left(2\omega t-2m\phi\right)}\right)+{\rm c.c.}\right)\nonumber\\ +\left(A^+_1A^-_1\hat{\boldsymbol{\mathcal{F}}}_2^{A^+A^-} {\rm e}^{\text{i}2\omega t}+A^+_1A^{-^*}_1\hat{\boldsymbol{\mathcal{F}}}_2^{A^+A^{-^*}} {\rm e}^{\text{i}2m\phi}+{\rm c.c.}\right). \end{gather}$$

All terms contributing to the forcing vector $\boldsymbol {\mathcal {F}}_2$ were extracted using symbolic calculus in Wolfram Mathematica (see supplementary material). The first-order solution is made of four different contributions of amplitude $A^+_1$, $A^{+^*}_1$, $A^-_1$ and $A^{-^*}_1$, and therefore it generates 10 different second-order forcing terms, $\hat {\boldsymbol {\mathcal {F}}}_2^{ij} \exp ({\text {i}(\omega ^{ij} t+m^{ij}\phi )})$, which exhibit a certain frequency and spatial periodicity, gathered in table 2. The two additional terms, $\hat {\boldsymbol {\mathcal {F}}}_2^{\hat {F}}$, appearing in the forcing expression (5.8), come from the spatially uniform axisymmetric external forcing typical of Faraday waves, whose amplitude was assumed to be of order $\epsilon ^2$.

Table 2.

Second-order nonlinear forcing terms gathered by their amplitude dependency, and corresponding azimuthal and temporal periodicity $(m^{ij},\omega ^{ij})$. Seven terms have been omitted as they are the complex conjugates.

All these forcing terms are non-resonant, as their oscillation frequencies and their spatial symmetries, through the azimuthal wavenumber, differ from those of the leading-order solution (see table 2). Hence no solvability conditions are required at the present order (Meliga et al. Reference Meliga, Chomaz and Sipp2009). We can thus look for a second-order solution as the superposition of the second-order response to the external forcing, $\hat {\boldsymbol {q}}_2^{\hat {F}}$, and 10 responses $\hat {\boldsymbol {q}}_2^{ij}$ to each single forcing term,

(5.9)$$\begin{gather} \boldsymbol{q}_2=|A^+_1|^2\hat{\boldsymbol{q}}_2^{A^+A^{+^*}}+|A^-_1|^2\hat{\boldsymbol{q}}_2^{A^-A^{-^*}}+\left(\hat{F}\hat{\boldsymbol{q}}_2^{\hat{F}} \exp\left({\text{i}\left(2\omega t+\hat{\varLambda}T\right)}\right)+{\rm c.c.}\right)\nonumber\\ +\left(A^{+^2}_1\hat{\boldsymbol{q}}_2^{A^{+^2}} \exp\left({\text{i}\left(2\omega t+2m\phi\right)}\right)+A^{-^2}_1\hat{\boldsymbol{q}}_2^{A^{-^2}} \exp\left({\text{i}\left(2\omega t-2m\phi\right)}\right)+{\rm c.c.}\right)\nonumber\\ +\left(A^+_1A^-_1\hat{\boldsymbol{q}}_2^{A^+A^-} {\rm e}^{\text{i}2\omega t}+A^+_1A^{-^*}_1\hat{\boldsymbol{q}}_2^{A^+A^{-^*}} {\rm e}^{\text{i}2m\phi}+{\rm c.c.}\right), \end{gather}$$

(the same form is assumed for $\eta _2$) each of which is computed as a solution of a linear forced problem

(5.10)\begin{equation} \left(\text{i}\omega^{ij}{{\boldsymbol{\mathsf{B}}}}-\tilde{{{\boldsymbol{\mathsf{A}}}}}_{m^{ij}}\right)\hat{\boldsymbol{q}}_2^{ij}=\hat{\boldsymbol{\mathcal{F}}}_2^{ij}, \end{equation}

with $m^{ij}$ and $\omega ^{ij}$ for $(i,j)$ from table 2 and which can be inverted (non-singular operator) so long as any of the combinations $(m^{ij},\omega ^{ij})$ is not an eigenvalue (none of them has $m^{ij}=\pm m$). As an example, the $\epsilon$-order eigensurface and some of the various second-order surfaces are shown in figure 2 for three different waves, i.e. $(m,n)=(1,2)$, $(3,2)$ and $(0,2)$. Owing to the symmetries of the system (given in (4.13)), some of the second-order responses corresponding to the generic $(m,n)$ wave have the same solution with opposite azimuthal velocity, therefore in figure 2 we show only the solutions with different surface shapes. Furthermore, as can be deduced from figure 2(m–r), in the axisymmetric case $(0,n)$ all the responses are axisymmetric with zero azimuthal velocity, thus some of the second-order responses share exactly the same solution. In this case, indeed, the second-order solution could be formulated a priori as the sum of three terms only, whose amplitudes are proportional to $\hat {F}$, $A^2_1$ (second harmonic) and $|A_1|^2$ (mean flow correction), respectively.

Figure 2.

(a–f) Upper subpanel: real part of the free surface elevation, Re$(\hat {\eta })$ associated with (a) mode $(1,2)$ and with (b–f) some of the corresponding second-order responses for different values of the static contact angle, $\theta _s$. The $\epsilon$-order solution is normalized such that the phase of the interface at the contact line in $\phi =0$ is zero and the corresponding slope is one, i.e. $\hat {\boldsymbol {q}}_1\rightarrow \hat {\boldsymbol {q}}_1 \exp ({-\text {i}\,\text {arctan}[\hat {\eta }_1(r=1,0)]})/(\partial \hat {\eta }_1(r,0)/\partial r|_{r=1})$. Lower subpanel: free surface visualization in terms of absolute value of the real part of the interface slope at $\theta _s=45^{\circ }$. The colourmaps were individually saturated for visualization purposes only. (g–l) Same as (a–f), but for mode $(3,2)$. (m–r) Same as (a–f), but for the axisymmetric mode $(0,2)$. Parameter setting: $R=0.035$ m; $h=0.022$ m; $\rho =997$ kg m$^{-3}$; $\mu =0.001$ kg m$^{-1}$ s$^{-1}$; $\gamma =0.072$ N m$^{-1}$; for which $Bo=166.2$ and $Re=20\,437$, and a static contact angle $\theta _s=45^{\circ }$. The light red boxes highlights the second-order response to the external forcing, i.e. second-order harmonic MW.

Of particular interest is the second-order response to the external forcing, whose interface shape is highlighted by the red boxes in figure 2. With the present scaling, the forcing enters at second order in the $z$-component of the momentum equation (see (2.1)). If the initial static interface is assumed to be flat ($\theta _s=90^{\circ }$), then the response $(\hat {\boldsymbol {q}}_2^{\hat {F}},\hat {\eta }_2^{\hat {F}})$, translates into a harmonic hydrostatic pressure modulation only, with a free surface remaining flat, i.e. $\hat {\boldsymbol {u}}_2^{\hat {F}}=\boldsymbol {0}$ and $\hat {\eta }_2^{\hat {F}}=0$, a case classically analysed in the literature. On the other hand, as shown in figure 2(c,i,o), if a static contact angle $\theta _s\ne 90^{\circ }$ is considered, then the $\epsilon ^0$-order static meniscus induces at order $\epsilon ^2$ axisymmetric meniscus capillary waves travelling from the sidewall to the interior and reflected back, which oscillates harmonically with the external forcing and with an amplitude proportional to the external forcing amplitude. In the present WNL analysis, these MW, which appear as concentric ripples (see figure 2c,i,o), as typically observed in experiments (Batson et al. Reference Batson, Zoueshtiagh and Narayanan2013; Shao et al. Reference Shao, Gabbard, Bostwick and Saylor2021a,Reference Shao, Wilson, Saylor and Bostwickb), will couple at third order with the first-order solution and will contribute to modify both the linear stability boundaries associated with the subharmonic Faraday tongues as well as the bifurcation diagram, i.e. wave amplitude saturation to finite amplitude. Furthermore, figure 2 clearly shows that a static contact angle $\theta _s\ne 90$, depending on its value (here only values of $\theta _s<90^{\circ }$ have been considered), modifies not only the damping coefficients and frequencies of the leading-order wave (see also Appendix A), but also its spatial shape and, as consequence, all the associated second-order responses, whose modifications may have a significant influence on the corresponding saturation to a finite amplitude.

5.5. Order $\epsilon ^3$: amplitude equation for standing waves

Lastly, at the $\epsilon ^3$-order we derive an amplitude equation for standing waves with a pinned contact line accounting for WNL modifications of the subharmonic Faraday threshold due to contact angle effects. The problem at order $\epsilon ^3$ is similar to the one obtained at order $\epsilon ^2$, as it appears as a linear system,

(5.11)\begin{equation} \left({{\boldsymbol{\mathsf{B}}}}\partial_t-\tilde{{{\boldsymbol{\mathsf{A}}}}}_m\right)\boldsymbol{q}_3=\boldsymbol{\mathcal{F}}_3, \end{equation}

forced by combinations of the previous-order solutions encompassed in $\boldsymbol {\mathcal {F}}_3$, that contains several nonlinear terms of various space and time periodicities and which we denote as $\hat {\boldsymbol {\mathcal {F}}}_3^{ij} \exp ({\text {i}(\omega t + m\phi )})$. Since many of these terms are resonant, as standard in multiple scale analysis, in order to avoid secular terms and solve the expansion procedure at the third order, a compatibility condition must be enforced through the Fredholm alternative (Friedrichs Reference Friedrichs2012). Such a compatibility condition imposes the amplitudes $A^+_1$ and $A^-_1$ to obey the following relation:

(5.12)\begin{equation} \frac{{\rm d}A^{{\pm}}}{{\rm d}t}={-}\sigma A^{{\pm}} + \zeta FA^{{\mp}^*} {\rm e}^{\text{i}\varLambda t / 2} + \chi_1 |A^{{\pm}}|^2A^{{\pm}} + \chi_2 |A^{{\mp}}|^2A^{{\pm}}, \end{equation}

where the physical time $t=T/\epsilon ^2$ has been reintroduced and where $\sigma =\epsilon ^2\hat {\sigma }$, $F=F_d/g=\epsilon ^2\hat {F}$ and $\varLambda =\epsilon ^2\hat {\varLambda }$. By considering the expansion $\boldsymbol {q}=\boldsymbol {q}_0+\epsilon A_1\hat {\boldsymbol {q}}_1\dots$, the small parameter $\epsilon$ is eliminated by defining the amplitude $A=\epsilon A_1$, so that everything is recast in terms of actual physical quantities (Bongarzone et al. Reference Bongarzone, Bertsch, Renaud and Gallaire2021a; Bongarzone, Guido & Gallaire Reference Bongarzone, Guido and Gallaire2022). The various coefficients are computed as scalar products between the adjoint global modes and the resonant forcing terms $\hat {\boldsymbol {\mathcal {F}}}_3^{ij}$, whose analytically complex expressions have been extracted from the third-order forcing using the symbolic calculus tools of Wolfram Mathematica. For instance, the complex coefficient $\zeta$ is evaluated as

(5.13)\begin{equation} \zeta=\frac{\int_{{V}}\hat{\boldsymbol{u}}_1^{{{\dagger}}^* A^{+}}\boldsymbol{\cdot}\hat{\boldsymbol{\mathcal{F}}}_{3,{NS}}^{\hat{F}A^{-^*}}\,r\,\text{d}r\text{d}z+\int_{\eta_0}\hat{\boldsymbol{u}}_1^{{{\dagger}}^* A^{+}}\boldsymbol{\cdot}\hat{\boldsymbol{\mathcal{F}}}_{3,{D}}^{\hat{F}A^{-^*}}\,r\,\text{d}r +\int_{\eta_0}\xi^{{{\dagger}}^* A^+}\hat{\mathcal{F}}_{3,{K}}^{\hat{F}A^{-^*}}\,r\,\text{d}r}{\int_{{V}}\hat{\boldsymbol{u}}_1^{{{\dagger}}^* A^{+}}\boldsymbol{\cdot}\hat{\boldsymbol{u}}_1^{A^{+}}\,r\,\text{d}r\,\text{d}z+\int_{\eta_0}\xi^{{{\dagger}}^* A^+}\hat{\eta}_1^{A^+}\,r\,\text{d}r}, \end{equation}

where ${V}$ denotes the fluid bulk domain, the dagger symbol refers to the adjoint eigenmode,

(5.14)\begin{equation} \xi = \left[-\frac{\eta_{0,r}}{Re}\left(\frac{\partial \hat{u}_{1z}}{\partial r}+\frac{\partial \hat{u}_{1r}}{\partial z}\right)+\left(-\hat{p}_1+\frac{2}{Re}\frac{\partial \hat{u}_{1z}}{\partial z}\right)\right] \end{equation}

(see also Viola & Gallaire Reference Viola and Gallaire2018) and the subscripts $_{NS}$, $_{D}$ and $_{K}$ designate the forcing components of $\hat {\boldsymbol {\mathcal {F}}}_3^{\hat {F}A^{-^*}}$ appearing in the $\epsilon ^3$-order Navier–Stokes equations, dynamic boundary condition and kinematic boundary condition, respectively. Analogous expressions hold for $\chi _1$ and $\chi _2$ by replacing $\hat {\boldsymbol {\mathcal {F}}}_3^{\hat {F}A^{-^*}}$ with $\hat {\boldsymbol {\mathcal {F}}}_3^{A^+A^{+^*}A^+}$ and $\hat {\boldsymbol {\mathcal {F}}}_3^{A^-A^{-^*}A^+}$, respectively. We notice that the adjoint eigenvector appearing in (5.13) does not need to be independently calculated. Indeed, Viola & Gallaire (Reference Viola and Gallaire2018) demonstrated that the linear operator ${{\boldsymbol{\mathsf{B}}}}$ and ${{\boldsymbol{\mathsf{A}}}}_m$ (the same applies to the shifted operator $\tilde {{{\boldsymbol{\mathsf{A}}}}}_m$) are self-adjoint, i.e. ${{\boldsymbol{\mathsf{B}}}}^{{\dagger} }={{\boldsymbol{\mathsf{B}}}}$ and ${{\boldsymbol{\mathsf{A}}}}_m^{{\dagger} }={{\boldsymbol{\mathsf{A}}}}_m$, with the adjoint eigenvalue being the complex conjugate of the direct one, $\lambda ^{{\dagger} }=\lambda ^*$. Then, from (4.13), (4.14) and (4.15), it follows that for the couple $(m,-\sigma +\text {i}\omega )$ associated with a direct mode, we have the relation

(5.15)\begin{equation} \left(\hat{u}_{1r}^*,-\hat{u}_{1\phi}^*,\hat{u}_{1z}^*,\hat{p}_1^*,\hat{\eta}_1^*\right)\rightarrow\left(\hat{u}_{1r}^{{{\dagger}}},\hat{u}_{1\phi}^{{{\dagger}}},\hat{u}_{1z}^{{{\dagger}}},\hat{p}_1^{{{\dagger}}},\hat{\eta}_1^{{{\dagger}}}\right), \end{equation}

which directly provides the desired adjoint mode without any further calculation. We also underline that due to the symmetry of the solution, the same value of $\zeta$ is obtained if one makes use of the scalar product between the adjoint mode for $A^-_1$ and the forcing term $\hat {\boldsymbol {\mathcal {F}}}_3^{\hat {F}A^{+^*}}$ (same for $\chi _1$ and $\chi _2$).

As anticipated before, the standing wave solution corresponds to the superposition of two balanced counter-rotating waves of same amplitude $A^+=A^-=A$. It follows that system (5.12) reduces to the single amplitude equation

(5.16)\begin{equation} \frac{{\rm d}B}{{\rm d}t}={-}\left(\sigma+\text{i}\varLambda/2\right) B + \zeta FB^* + \chi |B|^2B , \end{equation}

where the change of variable $A=B^{\text {i}\varLambda /2}$ has been introduced and where the complex coefficient $\chi$ is taken as the sum of $\chi _1$ and $\chi _2$. The form of (5.16) is totally equivalent to the normal form (1.1) postulated by Douady (Reference Douady1990) using symmetry arguments only. Its structure indeed does not depend on the boundary conditions or on the mode shape, nevertheless its coefficients do. In the present work these complex coefficients, $\zeta$ and $\chi$, as well as the frequency and damping of the wave, $\omega$ and $\sigma$, are formally computed by taking into account the full hydrodynamic system, whose solution is exact at numerical convergence. The damping coefficient must be small enough, but its value is numerically computed, rather than estimated heuristically. Most importantly, $\zeta$ and $\chi$, through the WNL formulation presented above, encompass in a formal manner, although within the assumptions of validity of a single-mode WNL theory, the effect of the static contact angle and of the coupling with harmonic MW on the subharmonic Faraday threshold of standing viscous capillary–gravity waves with pinned contact line.

5.6. Linear stability of the amplitude equation: subharmonic Faraday tongues

Here we perform the stability analysis of the amplitude equation (5.16), which prescribes the marginal stability boundaries, typically known as Faraday tongues. By turning to polar coordinates

(5.17)\begin{equation} B=|B|{\rm e}^{\text{i}\varPhi},\quad -\left(\sigma+\text{i}\varLambda/2\right)=c_1 {\rm e}^{\text{i}\varphi_1},\quad \zeta=c_2 {\rm e}^{\text{i}\varphi_2},\quad \chi=c_3 {\rm e}^{\text{i}\varphi_3}, \end{equation}

splitting the modulus and phase parts of (5.16) and introducing the change of variable $\varTheta =\varPhi -\varphi _2/2$, we obtain the following system:

(5.18)\begin{gather} \frac{{\rm d}|B|}{{\rm d}t}=c_1\cos{\left(\varphi_1\right)}|B|+c_2\cos{\left(2\varTheta\right)}F|B|+c_3\cos{\left(\varphi_3\right)}|B|^3, \end{gather}

(5.19)\begin{gather} \frac{{\rm d}\varTheta}{{\rm d}t}=c_1\sin{\left(\varphi_1\right)}-c_2\sin{\left(2\varTheta\right)}F+c_3\sin{\left(\varphi_3\right)}|B|^2. \end{gather}

Equation (5.18) admits two possible equilibria $(d/dt=0)$, having $|B|=0$ and $|B|\ne 0$, respectively. We first focus on the stability of the trivial stationary solution, $|B|=0$. By eliminating $\varTheta$ from (5.18)–(5.19), the linear threshold or marginal stability boundaries (subharmonic Faraday tongues) are readily obtained (Douady Reference Douady1990; Rajchenbach & Clamond Reference Rajchenbach and Clamond2015),

(5.20)\begin{equation} F_{th}^L=\left(F_d/g\right)_{th}^L=c_1/c_2\quad \longrightarrow\quad F_{th}^L={\pm}|\zeta|^{{-}1}\sqrt{\sigma^2+\left(\varOmega_d/2-\omega\right)^2}, \end{equation}

where the relation $\varLambda =\varOmega _d-2\omega$ has been reintroduced and which predicts the lowest threshold, $F_{th,min}^L=\sigma /|\zeta |$, at $\varOmega _d=2\omega$. The forcing amplitude at which the instability appears is therefore proportional to its dissipation, $\sim \sigma$ (note that this is true only for subharmonic resonances, e.g. the threshold for harmonic tongues is expected to scale as $\sim \sigma ^{1/2}$, see Rajchenbach & Clamond (Reference Rajchenbach and Clamond2015)). Moreover, $F_{th}^L$ depends on the coefficient $\zeta$, which is produced by the interaction of the first-order response, proportional to the amplitude $A^{\pm }$, with the second-order response to the external forcing, proportional to $\hat {F}$. Therefore, contact angle modifications of the leading-order solution and harmonic MW (see figure 2) enter directly in the calculation of $\zeta$, whose value contributes to the definition of the marginal stability boundaries. Presence of a static meniscus, as widely discussed in § 4, also modifies the natural frequency $\omega$ and the damping $\sigma$. Lastly, it is important to note that within the subharmonic tongue (linearly unstable) the signal is periodic with a frequency equal to half the driving frequency, $\varOmega _d/2$, whereas out of the tongue (linearly stable) it is periodic with frequency $\varOmega _d$, due to the presence of harmonic meniscus waves.

5.6.1. Brimful condition: validation with the inviscid analysis by K13 for $\theta _s=90^{\circ }$

The most comprehensive investigation of Faraday thresholds with pinned contact line that the authors are aware of is that of K13 (see table 1), who considered the case of a perfect brimful condition (meniscus-free) in the inviscid limit. Unlike the classic case of an ideal moving contact line, K13 showed that the pinned contact line problem can be recast into an infinite system of coupled Mathieu equations taking the following form:

(5.21)\begin{equation} \frac{{\rm d}^2\boldsymbol{y}}{{\rm d}\tau^2}+\left({{\boldsymbol{\mathsf{P}}}}-2{{\boldsymbol{\mathsf{Q}}}}\cos{2\tau}\right)\boldsymbol{y}=\boldsymbol{0}, \end{equation}

where matrices ${{\boldsymbol{\mathsf{P}}}}$ and ${{\boldsymbol{\mathsf{Q}}}}$, obtained via projection onto the test function space, are in general not diagonal (for a free contact line ${{\boldsymbol{\mathsf{P}}}}$ and ${{\boldsymbol{\mathsf{Q}}}}$ are diagonal, so that (5.21) reduces to (2.14) of Benjamin & Ursell (Reference Benjamin and Ursell1954), i.e. uncoupled Mathieu equations). Three different methods (Nayfeh & Mook Reference Nayfeh and Mook1995) can be used to solve (5.21) , namely, (i) the mapping at a period (given by the Floquet theory), (ii) Hill's infinite determinant method (used by Kumar & Tuckerman (Reference Kumar and Tuckerman1994)) and (iii) the multiple scale method. The first two techniques were used in K13 and, particularly, the first one was employed in order to describe the so-called combination resonance tongues (indicated by the black arrows in figure 3), which are not studied in the present work which is focused on subharmonic tongues only (see K13 for a thorough discussion). The disadvantage of the multiple scale method is that generally it is not suitable for the exploration of a large part of the parameter space, however, as anticipated in the introduction, the application of the first two techniques is challenging when the initial free surface is not assumed to be flat. Here we use the inviscid results provided by K13 to validate the present WNL model for prediction of subharmonic instability onset in the limit of high Reynolds numbers (e.g. $Re$ is assumed to be $\sim 10^6$ in the present viscous analysis).

Figure 3.

Inviscid stability plots associated with modes $(1,1)$ and $(1,2)$ for two different Bond numbers, i.e. $Bo=1000$ and $100$, and for a depth $h/R=H=1$. The grey shaded regions have not been reproduced in this work, but rather they have been simply taken from figure 10 of K13. subharmonic tongues are denoted by the subscript $_{sh}$. For computational reasons, the instability regions (grey shaded) were obtained in K13 by truncating the number of basis function $N_{K13}$ to 2, although convergence of the natural frequencies was achieved by taking $N_{K13}=30$, as stated by K13 in his table 1 (with a systematic underestimation of approximately 5 %). The vertical black dash–dot lines correspond to the converged results reported in table 1 of K13. The blue solid lines correspond to the present numerical prediction computed through (5.20) for $Re=10^6$, while the coloured dash–dot lines denote the present Faraday tongues shifted by 5 %. Black and coloured lines have been added on top of the original figure from K13.

A quantitative comparison of the prediction of subharmonic Faraday thresholds with results by K13 is shown in figure 3 for $\theta _s=90^{\circ }$, $h/R=H=1$, for two different Bond numbers and for two non-axisymmetric modes, i.e. $(1,1)$ and $(1,2)$. For computational reasons, the instability regions (grey shaded) computed by K13 were obtained by truncating the number of basis function $N_{K13}$ to 2, although convergence of the natural frequencies was achieved by taking $N_{K13}=30$, as stated by K13 in his table 1, causing a systematic underestimation of approximately 5 %. The vertical black dash–dot lines, corresponding to the converged natural frequencies reported in table 1 for $N_{K13}=30$, agrees perfectly with the present prediction, which prescribes the correct slope of the right-hand and left-hand marginal stability boundaries (blue solid lines). If the present prediction is shifted by $-5\,\%$ (orange dash–dot lines), the results match. Hence we can conclude that the present model is congruent with the analysis by K13 and it prescribes correctly the subharmonic Faraday tongues for a pinned contact line case in the limit of validity of the WNL model, i.e. small external forcing amplitude and small detuning.

5.6.2. Brimful condition: comparison with recent experiments by S21 for $\theta _s=90^{\circ }$

From the knowledge of the authors, no systematic calculations of the linear subharmonic Faraday tongues for pinned contact line and including viscous dissipation are reported in the literature. With regard to small circular cylinder experiments, this configuration was recently studied by Shao et al. (Reference Shao, Wilson, Saylor and Bostwick2021b) (S21). By properly filling the container they could reproduce an initially flat static-free surface, which remains stable and flat below the Faraday threshold and thereby they could derive experimentally the boundaries of the unstable regions. Their experimental measurements (extracted from figure 4 of S21) are illustrated in figure 4(a), as coloured filled circles, together with our numerical prediction from (5.20) (coloured solid lines). Shao et al. (Reference Shao, Wilson, Saylor and Bostwick2021b) also employed a Rayleigh–Ritz approach (Bostwick & Steen Reference Bostwick and Steen2009) to estimate numerically the natural frequency in the inviscid limit and this result, which showed a good agreement with their experiments, is reported for completeness in figure 4(a) as vertical black dash–dot lines.

Figure 4.

(a) Coloured solid lines: boundaries of the subharmonic Faraday tongues predicted by (5.20) in the forcing acceleration amplitude-forcing frequency dimensional space, $(f_d,F_d)$. Here the static contact angle was set to $\theta _s=90^{\circ }$. The coloured filled circles corresponds to the original experimental values extracted from figure 4 of S21 for different waves $(m,n)$. The black dash–dot lines correspond to their inviscid numerical calculation. Parameters: $R=0.034925$ m; $h=0.022$ m; $\rho =1000$ kg m$^{-3}$; $\mu =0.001$ kg m$^{-1}$ s$^{-1}$; and $\gamma =0.072$ N m$^{-1}$; for which $Bo=165.5$ and $Re=20\,371$. Coloured bands: marginal stability boundaries computed for a container radius $R=(0.034925-0.000254)$ m (right boundary) and $R=(0.034925+0.000254)$ m (left boundary). (b) Modification of the linearly unstable regions due to contact angle effects, where the results for three values of $\theta _s$, including $90^{\circ }$ (black dotted lines) as in (a), are compared for a nominal radius $R=0.035$ m.

The present numerical analysis for $\theta _s=90^{\circ }$ predicts the occurrence of the same subharmonic single-mode instabilities in the selected frequency window. In agreement with experimental observations, the viscous WNL analysis prescribes a minimum onset acceleration that is nearly constant for all $(m,n)$-modes in the range $f_d\in [10,20]$ Hz with a discrete spectrum of subharmonic resonances. Moreover, the WNL model predicts correctly the coefficient $\zeta$, which prescribes the slope of the transition curves for all tongues.

All the experimental frequencies are slightly larger than the ones predicted here and this shift is roughly of the order of +1 % for all measurements. It is difficult to attribute a positive 1 % shift to a specific cause, especially because the pinned contact line configuration is known to produce the largest frequencies among the possible contact line boundary conditions, e.g. a free contact line. Presence of free surface contamination (surface film) is expected to slightly increase the rigidity of the free surface, leading to higher resonance frequencies, but also to larger damping coefficients, which reduce the frequencies (Miles Reference Miles1967; Henderson & Miles Reference Henderson and Miles1990, Reference Henderson and Miles1994). However, any evidence of surface contamination is reported in S21. In the present case, such a slight systematic mismatch is more likely to be caused by small incongruities between numerics and experiments. For instance, in this case the Bond number is relatively low, $Bo=165.5$, so that small variations in the value of the surface tension or, alternatively, geometrical tolerances on the container radius could contribute to shift the tongues slightly.

In S21 the authors report the nominal container radius $R=0.035$. We have written to the authors, who have kindly provided us with the technical drawing of their cylindrical container. The actual nominal (inner) radius is $R=0.034925\,$m. Unfortunately, the tolerance on the inner radius is not specified. Nevertheless, given the tolerances specified by the manufacturer on the outer radius, i.e. $\pm 0.000254\,$m, it is natural to assume at least the same value for the inner one. In figure 4(a) the subharmonic tongues computed for the nominal radius are shown as coloured solid lines, whereas the coloured bands are associated with the geometrical tolerance on the container radius. One can see that a value of $R=0.034925-0.000254=0.034671\,$m (right-hand boundaries) is sufficient to produce a +1 % frequency shift so to achieve a fairly good agreement with the experimental measurements.

For completeness, the values of the damping coefficients, natural frequencies and of the normal form coefficient $\zeta$ for two different static contact angles used in figure 4 are given in table 3.

Table 3.

Non-dimensional natural frequencies, damping coefficients ($\lambda$ is the eigenvalue $\lambda =-\sigma +\text {i}\omega$) and complex normal form coefficient $\zeta =\zeta _{\text {R}}+\text {i}\zeta _{\text {I}}$ for both $\theta _s=90^{\circ }$ and $\theta _s=45^{\circ }$, associated with the modes shown in figure 4 and computed for $R=0.034925$ m, $h=0.022$ m, $\rho =1000$ kg m$^{-3}$, $\mu =0.001$ kg m$^{-1}$ s$^{-1}$ and $\gamma =0.072$ N m$^{-1}$, for which $Bo=165.5$ and $Re=20\,371$. The number of points in the radial and axial directions for the GLC grid used is this calculation is $N_r=N_z=80$, for which convergence up to the third digit is achieved.

5.7. Weakly nonlinear threshold and bifurcation diagram

In this paragraph, we focus on the stability of the non-trivial equilibrium, $|B|\ne 0$, of system (5.18)–(5.19). Again, for stationary solutions, we find by eliminating $\varTheta$ that

(5.22)\begin{equation} c_3|B|^2={-}c_1\cos{\left(\varphi_1-\varphi_3\right)}\pm\sqrt{c_2^2F^2-c_1^2\sin^2{\left(\varphi_1-\varphi_3\right)}}, \end{equation}

with physical real solutions for $F\ge \frac {c_1}{c_2}|\sin {(\varphi _1-\varphi _3)}|$. This well known result prescribes either a supercritical or a subcritical transition when the marginal stability boundaries are crossed, i.e. by changing forcing frequency and amplitude. The location of the hysteresis depends on the sign of the nonlinear coefficient $\chi$ (Kovacic, Rand & Sah Reference Kovacic, Rand and Sah2018), which assumes the meaning of a nonlinear detuning, while the boundary of the hysteresis region in the parameter space is defined by the nonlinear threshold

(5.23)\begin{equation} F_{th}^{NL}=c_1/c_2|\sin{\left(\varphi_1-\varphi_2\right)}|. \end{equation}

In figure 5 the nonlinear wave amplitude saturation, for a fixed external acceleration amplitude, $F_d$, and for a varying excitation frequency, $\varOmega _d$, is shown for two different modes, $(0,2)$ and $(3,2)$, and for different static contact angle values. The linear acceleration threshold (Faraday tongue) is plotted versus a normalized driving frequency in order to better compare the difference between the two cases with $\theta _s=90^{\circ }$ (flat static surface, brimful condition) and $45^{\circ }$ (static meniscus and MW, nearly brimful condition). As previously discussed, contact angle modifications on the linear thresholds are only weak. When a concave ($\theta _s<90^{\circ }$) static meniscus is considered, the damping is generally higher, the shape of the mode is, however, modified, leading to a slightly different value of the complex linear coefficient $\zeta$ (see table 3), which also encompasses the second-order coupling between parametric and MW. As a consequence, the minimum onset acceleration, given by the ratio $\sigma /|\zeta |$, is often comparable.

Figure 5.

Linear acceleration threshold (Faraday tongue) (left-hand $y$-axis; thin solid lines) and saturated wave amplitude, $|B|$, (right-hand $y$-axis; thick solid lines) for a fixed acceleration amplitude $F_d=0.5$ m s$^{-2}$, while the driving frequency is varied. Stable branches for $|B|$ are shown as solid lines, while unstable branches as dashed lines. Two different modes corresponding, namely (a) $(m,n)=(3,2)$ and (b) $(0,2)$, are shown. Different static contact angle are considered. The frequency is normalized with twice the natural frequency of the corresponding excited mode, so that the lowest linear threshold occurs for $\varOmega _d/2\omega =1$ for all $\theta _s$. At convergence (GLC grid $N_r=N_z=80$), the complex nonlinear amplitude equation coefficient, $\chi =\chi _R+\text {i}\,\chi _I$, for mode $(0,2)$ (inset in panel (b)), assumed the values, $\chi ^{90^{\circ }}=-0.0909-\text {i}\,1.9094$ and $\chi ^{45^{\circ }}=-0.0184-\text {i}\,0.5617$. Geometrical and physical parameters are set as in figure 2.

Supercritical and subcritical bifurcations of Faraday waves have been widely discussed in the literature (see, for instance, Douady (Reference Douady1990) and Rajchenbach & Clamond (Reference Rajchenbach and Clamond2015), among other references), hence we limit here to recall that if $\cos {(\varphi _1-\varphi _3)}>0$, or alternatively $\varLambda =\varOmega _d-2\omega >-2\sigma \chi _R/\chi _I$, then the bifurcation is supercritical, while if $\cos {(\varphi _1-\varphi _3)}<0$, or $\varLambda =\varOmega _d-2\omega <-2\sigma \chi _R/\chi _I$, the transition is subcritical, the sign of $\chi _I$ determines whether hysteresis occurs on the left-hand side or on the right-hand side. The inferior boundary of the hysteresis region in the $(\varOmega _d,F_d)$-plane is defined by (5.23). In other words, the ratio $\chi _R/\chi _I$, through the relation $\varphi _3=\tan ^{-1}{(\chi _I/\chi _R)}$, determines the importance of the subcritical region in the parameter space (Hsu Reference Hsu1977; Gu & Sethna Reference Gu and Sethna1987; Meron Reference Meron1987; Nayfeh & Mook Reference Nayfeh and Mook1995; Douady Reference Douady1990).

We underline that the amplitude equation coefficients setting the nonlinear threshold and the bifurcation diagram are not calibrated from experimental data, but their values are here computed numerically from first principles through our WNL analysis.

5.7.2. The imperfect bifurcation diagram: tailing effect

As shown in figure 5, the linear threshold given by (5.20) prescribes a stable solution outside the subharmonic Faraday tongues (see figure 4) with a stationary mode amplitude $|B|=0$. Nevertheless, we remind the reader that the total solution, e.g. in terms of free surface elevation, is given by the sum of the solutions at the various orders in $\epsilon$, i.e. $\eta =\eta _0+\eta _1(B)+\eta _2(F_d,B^2,|B|^2)$. In particular, MW, whose amplitude is proportional to the external acceleration amplitude, $F_d$, are contained in the second-order response $\eta _2$. If one considers an axisymmetric dynamics, e.g. $(0,2)$, the amplitude of the centreline elevation is a suitable quantity to monitor the free surface stability and thus to depict a comprehensive bifurcation diagram. This is done in figure 6, where such a bifurcation diagram for $(0,2)$ is reported for different excitation angular frequencies in a range which gathers both supercritical and subcritical bifurcations. Figure 6 clearly shows that, when a nearly brimful condition is considered, e.g. $\theta _s<90^{\circ }$, the subharmonic parametric waves, stable outside the Faraday tongues, do not bifurcate from the rest state (as for $\theta _s=90^{\circ }$), but rather from the MW solution ($\propto F_d$), oscillating harmonically with the driving frequency. This produces a so-called imperfect bifurcation diagram, which displays a tailing effect (highlighted by the black thin solid line) (Virnig et al. Reference Virnig, Berman and Sethna1988). The bifurcation diagram of figure 6 is also reminiscent of that presented by Batson et al. (Reference Batson, Zoueshtiagh and Narayanan2013), although they focus on harmonic parametric waves.

Figure 6.

Bifurcation diagram associated with $(m,n)=(0,2)$ (see also figure 5a) and for a static contact angle $\theta _s=45^{\circ }$. Here the dimensional centreline amplitude (axisymmetric dynamic) is reconstructed by summing the various-order solutions, i.e. $\eta =\eta _0+\eta _1+\eta _2$ and it is plotted versus the external forcing acceleration for a fixed excitation angular frequency, while different colours correspond to different forcing frequencies. The tailing effect (imperfect bifurcation diagram) produced by presence of harmonic MW and indicated by the black thin solid line (the amplitude of MW grows linearly with $F_d$, independently of the parameter combination $(\varOmega _d,F_d)$), is well visible in the right-hand inset. Coloured solid lines are used for stable branches, while coloured dashed lines for the unstable ones. The hysteretic loop is indicated by the green arrows. The centreline amplitude is simply computed as $\max _{t}\eta (r=0,t)/2-\min _{t}\eta (r=0,t)/2$.

6.1. Procedure

To start, the shape of the static meniscus, computed in MATLAB by solving (4.1) with its boundary conditions (prescribing a static contact angle value, e.g. $\theta _s=45^{\circ }$) was loaded in COMSOL Multiphysics and the static domain was meshed. First, simulations were initialized for time $t=0$ with a BDF1 scheme giving a zero velocity field and hydrostatic pressure $p=-z$ as initial conditions. A body forcing, corresponding to the non-dimensional time-dependent gravity acceleration, $-1+(F_d/g)\cos {\varOmega _d t}$, was assigned. The starting point of the grey arrows in figure 7(b) indicates the combination of external control parameter $(\varOmega _d,F_d)$ (coloured markers), chosen to initiate the simulations, as described above. Once the stationary state for these initial DNS was established, a continuation procedure (directions of the arrows), by slightly adjusting the external amplitude acceleration and angular frequency, was adopted in order to speed up the computations for all the other combinations of parameters here considered (see figure 7).

Figure 7.

(a) Faraday tongue (black solid line) for the axisymmetric mode $(0,2)$ and for a static contact angle $\theta _s=45^{\circ }$. Forcing frequency and amplitude in the $(\varOmega _d,F_d)$-space, corresponding to the DNS points in (b), are indicated by coloured filled markers. Note that the frequency in the $x$-axis is normalized using the natural frequency $\omega =3.16$ computed for $\theta _s=45^{\circ }$. The grey arrows denote the direction followed in the continuation procedure for DNS. For completeness, the Faraday tongue for $\theta _s=90^{\circ }$ is reported as grey dashed line. (b) Associated bifurcation diagram: WNL prediction (lines) versus DNS (markers). The unstable branch is displayed as coloured dashed lines. The black solid line indicating the slop of the meniscus wave response is also given to guide the eyes. The centreline amplitude is computed as $\max _{t}\eta (r=0,t)/2-\min _{t}\eta (r=0,t)/2$.

6.2. Amplitude saturation and free surface reconstruction: WNL versus DNS

The WNL prediction (5.22) for the finite amplitude saturation is compared with DNS in figure 7. The selected combinations of control parameters, i.e. $(\varOmega _d,F_d)$, for DNS calculations are indicated by coloured markers in figure 7(a), where the grey arrows display the direction followed in the continuation procedure. Once the stationary state is established, i.e. the wave amplitude saturates, and the underlying dynamics being axisymmetric, the centreline free surface elevation is used as a reference measure of the free surface destabilization and of its saturation to finite amplitude. The DNS results are therefore compared with the WNL prediction, where the centreline dynamics is reconstructed by evaluating $\eta =\eta _0+\eta _1+\eta _2$ in $r=0$ for any time. The resulting amplitude comparison is shown in figure 7(b). At small forcing amplitude below the Faraday threshold (see also figure 7b), only harmonic travelling MW, whose amplitude is proportional to $F_d$, are observed in the DNS, consistently with the WNL model (straight line in figure 7b). In this small amplitude regime, the WNL model and DNS are in fairly good agreement in terms of free surface dynamics (see figure 8a,b). The frequency spectrum in figure 8(b) clearly highlights the harmonic nature of these zero-threhsold MW, directly forced by the container sidewalls as soon as the vertical excitation starts.

Figure 8.

Plot of WNL (black) versus DNS (red) below Faraday threshold (outside the Faraday tongue) for $\varOmega _d/\omega ^{45^{\circ }}=0.9804$ and $F_d=0.85$ m s$^{-2}$ (see figure 7). (a) Free surface shape computed when the centreline elevation is maximum. For completeness, the shape of the static meniscus for $\theta _s=45^{\circ }$ is reported as a black dotted line. (b) Corresponding frequency spectrum: power spectral density (PSD) versus the dimensional oscillation frequency of the system response.

By increasing the external acceleration amplitude $F_d$, the stability boundary (Faraday tongue in figure 7a) is eventually crossed and the parametric wave emerges on the top of edge waves, i.e. it bifurcates from the new stable and harmonically oscillating configuration.

Employing a continuation technique by progressively increasing/decreasing the forcing amplitude at different driving frequencies, several DNS were performed in both the supercritical and subcritical regime (filled coloured circles and triangles in figure 7, respectively). The agreement between DNS and WNL prediction in terms of amplitude saturation is found to be fairly good. Moreover, as figure 7(a) shows, DNS are consistent with the frequency shift caused by the presence of the static meniscus for $\theta _s=45^{\circ }$. As an example, the fully nonlinear free surface dynamics obtained from DNS for $\varOmega _d/\omega ^{45^{\circ }}=1.0054$ and $F_d=0.675$ m s$^{-2}$ is compared with the WNL reconstruction in figure 9(a–c) for three different time instants, while the corresponding centreline elevation and frequency spectrum are provided in figures 9(g) and 9(h), respectively.

Figure 9.

Plot of WNL (black) versus DNS (red) above Faraday threshold (within the Faraday tongue) for $\varOmega _d/\omega ^{45^{\circ }}=1.0054$ and $F_d=0.675$ m s$^{-2}$ (see figure 7). (a–c) Comparison in term of free surface reconstruction for three different time instants: (a) when the centreline elevation is maximum; (b) when it is zero and equal to the static meniscus position; and (c) when it is minimum. For completeness, the shape of the static meniscus for $\theta _s=45^{\circ }$ is reported as a black dotted line. (d–f) Full three-dimensional visualization extracted from the DNS. (g) Centreline elevation versus time associated with (a–c). Here $t_0$ is an arbitrary time instant. The constant value of the static meniscus elevation at $r=0$ is shown as a black dotted line. (h) Frequency spectrum computed from the time series shown in (g): PSD versus the dimensional oscillation frequency of the system response.

The WNL model is in agreement with the DNS, which consistently predicts the excitation of a dominant subharmonic parametric wave $(0,2)$, coupled with smaller amplitude harmonic meniscus waves as well as with higher-order harmonics (only second harmonics are included in the asymptotic expansion up to the third order in $\epsilon$).

As a final comment to this section, while not the purpose of the present analysis, a few DNS were performed at higher external acceleration amplitudes, in the parameter region far from the hypotheses of validity of the WNL theory. For the case of figure 7(b), preliminary observations revealed that DNS tends to diverge when the centreline elevation approaches a value of approximatively 5 mm, suggesting a potential transition to a highly nonlinear wave-breaking condition and eventually to a finite-time singularity with intense jet formation (Basak, Farsoiya & Dasgupta Reference Basak, Farsoiya and Dasgupta2021). See also Das & Hopfinger (Reference Das and Hopfinger2008) for a detailed investigation of the occurrence of such a phenomenon in Faraday experiments.